Truncated many-body dynamics of interacting bosons: A variational principle with error monitoring

# Truncated many-body dynamics of interacting bosons: A variational principle with error monitoring

by Kang-Soo Lee and Uwe R. Fischer

I apologize for not finishing this paper. I just open this in the internet. But this document is flawed as it is originally written in LaTeX and not translated to the docuK format perfectly yet. This paper is published in Int. J. Mod. Phys. B 84, 1550021 (2014) and arXiv>cond-mat>arXiv:1301.2199. ##[.no-sec-N] Abstract We introduce a scheme to describe the evolution of an interacting system of bosons, for which the field operator expansion is truncated after a finite number of modes, in a rigorously controlled manner. Using McLachlan's principle of least error , we find a set of equations for the many-body state. As a particular benefit, and in distinction to previously proposed approaches , our approach allows for the dynamical increase of the number of orbitals during the temporal evolution, because we can rigorously monitor the error made by increasing the truncation dimension. The additional orbitals, determined by the condition of least error of the truncated evolution relative to the exact one, are obtained from an initial trial state by a method we call steepest constrained descent. ## PH

• 2022-07-06: Update (images, bold, italic and so on.)
• 2014-12-17: To the docuK.
• 2014-06-23: First posting?

## TOC ## Introduction When we treat an atom, such as ^{87}Rb, ^{23}Na, and ^{7}Li which consists of an even [Z (number of neutrons and protons) +A (number of electrons)] number of fermions, as a whole, the Fock space creation and annihilation operators of these composites satisfy bosonic commutation relations. Since the experimental realization of Bose-Einstein condensates , a large variety of experiments with these bosonic isotopes have opened up a fascinating mesoscopic and macroscopic quantum world . After an initial period, concentrating on the effective single-particle physics of these ultracold dilute gases , more recently their many-body physics came into focus, revealing rich and hitherto unexpected possibilities to test fundamental correlation properties at the microscopic level .

For fermions, the commutator is replaced by the anticommutator \{\ \ , \ \ \}, : \begin{align*} \{a^{\,}_i, a^\dagger_j\} &\equiv a^{\,}_i a^\dagger_j + a^\dagger_j a^{\,}_i = \delta_{i j}, \\ \{ a^\dagger_i, a^\dagger_j \} &= \{ a^{\,}_i, a^{\,}_j \} = 0. \end{align*} Therefore, exchanging disjoint (i.e. i \ne j) operators in a product of creation or annihilation operators will reverse the sign in fermion systems, but not in boson systems Wiki :: Creation and annihilation operators :: Creation and annihilation operators in quantum field theories
and Wiki :: Second quantization
and Wiki :: Quantum field theory. But I am not currently understanding the reason why the composite of even fermions satisfy bosonic commutation relation \begin{align*} [ a^{\,}_i, a^\dagger_j ] &= a^{\,}_i a^\dagger_j - a^\dagger_j a^{\,}_i = \delta_{i j}, \\ [ a^\dagger_i, a^\dagger_j ] &= [ a^{\,}_i, a^{\,}_j ] = 0. \end{align*} I thought the composition of any particles in a set state \alpha, the creation and annihilation operators can be expressed like \begin{align*} z^\dagger_\alpha = \Big( a^\dagger_{1_\alpha} a^\dagger_{2_\alpha} b^\dagger_{3_\alpha} \cdots c^\dagger_{N_\alpha} \Big) \\ z^{\,}_\alpha = \Big( a^{\,}_{1_\alpha} a^{\,}_{2_\alpha} b^{\,}_{3_\alpha} \cdots c^{\,}_{N_\alpha} \Big) , \end{align*} where a^\dagger means creation of a particle, b^\dagger means creation of b particle, and so on. Since these a, b, \cdots c particles are fermions, I thought that the commutation relation of composite particle z breaks down. \begin{align*} [ z^{\,}_\alpha, z^\dagger_\beta ] &= z^{\,}_\alpha z^\dagger_\beta - z^\dagger_\beta z^{\,}_\alpha = \delta_{\alpha \beta}, \\ [ z^\dagger_\alpha, z^\dagger_\beta ] &= [ z^{\,}_\alpha, z^{\,}_\beta ] = 0. \end{align*} \begin{align*} [ z^{\,}_\alpha, z^\dagger_\beta ] &= \Big( a^{\,}_{1_\alpha} a^{\,}_{2_\alpha} b^{\,}_{3_\alpha} \cdots c^{\,}_{N_\alpha} \Big) \Big( a^\dagger_{1_\beta} a^\dagger_{2_\beta} b^\dagger_{3_\beta} \cdots c^\dagger_{N_\beta} \Big) - \Big( a^\dagger_{1_\beta} a^\dagger_{2_\beta} b^\dagger_{3_\beta} \cdots c^\dagger_{N_\beta} \Big) \Big( a^{\,}_{1_\alpha} a^{\,}_{2_\alpha} b^{\,}_{3_\alpha} \cdots c^{\,}_{N_\alpha} \Big) = \delta_{\alpha \beta}, \\ [ z^\dagger_\alpha, z^\dagger_\beta ] &= [ z^{\,}_\alpha, z^{\,}_\beta ] = 0. \end{align*} And I thought any more than double creation of the same state \alpha of z kills any many-body Fock state. z^\dagger_\alpha z^\dagger_\alpha | \text{any} \rangle = 0 , since the individual fermions are created twice on the same state. So treating a composite particle, which is composed of even fundamental fermion particles, as a boson seems to be wrong for me.

The extension to a true many-body physics, incorporating quantum correlations beyond mean-field, requires, however, vast computational resources when both the number of particles and the interaction between those particles increases. Therefore, a simplification of the problem by truncating the field operator expansion to a finite number of modes (or, as an equivalent term, single-particle orbitals) has been commonly utilized to obtain results relevant to the prediction of experiments in trapped bosonic quantum gases. The most extreme truncation, the semiclassical form of mean-field theory, retaining just one orbital, gives the well-known Gross-Pitaevski\check{i} equation. Without the aid of contemporary computers, it seemed to be inevitable until most recently, particularly in out-of-equilibrium situations far away from the ground state, to reduce the complexity of the problem at hand as much as possible, and hence to use the Gross-Pitaevski\check{i} equation approach. With the increased interest in many-body physics, however, there arose the necessity to go beyond the all-too-simplified mean field approach of the Gross-Pitaevski\check{i} equation. The accuracy of predictions on many-body correlations and the corresponding response functions will obviously increase with a less severe degree of truncation, though the solutions will not be exact still. #####/ Necessity of truncation of field operator expansion to render problems computable To derive the equations of many-body evolution, various variational approaches can be employed. Historically the first was the variational ansatz of Dirac and Frenkel , followed by McLachlan's variational principle and the time-dependent variational principle (TDVP), which is a principle of stationary action . Therefore, there are various, not necessarily equivalent, choices of variational principle for finding the equation of motion of the truncated many-body evolution. The Dirac-Frenkel principle imposes \langle\delta\Phi| \hat{H}-i\partial_t |\Phi\rangle = 0 (\hbar \equiv 1), where \langle\delta \Phi| denotes any possible variations of the many-body state \langle\Phi| with respect to a given set of variational parameters, whereas McLachlan's principle requires that the error of many-body evolution must be minimized. #####/ %\col{Sometimes you use \Psi, sometimes \Phi for the many-body state; is there supposed to be a difference, for example one is truncated the other exact? If not, we should use one symbol, e.g. \Phi and its variation.} On the other hand, the TDVP, as stated, requires stationarity of a given action. The three principles thus support quite different doctrines. Applying either the TDVP or the Dirac-Frenkel's principle, the authors in have proposed a method they called MCTDHB (Multi-Configurational Time-Dependent Hartree method for Bosons). This approach has, for example, provided tools for the description of the fragmentation of bosonic many-body states . We will describe below in detail that, besides its many beneficial properties, the MCTDHB method is incomplete in certain situations. Specifically, when the single-particle density matrix (SPDM) becomes singular, i.e. noninvertible, the method fails. As a consequence, MCTDHB does not provide a way to propagate, for example, a \emph{pure} single condensate into a fragmented condensate many-body state. Although MCTDHB provides an important tool to describe the many-body physics of interacting bosons, the method therefore lacks the possibility to directly connect the phenomena of condensation and fragmentation. #####/ %will show: ambiguity of [previously employed Dirac-Frenkel variational principle for many-body evolution %MacLachlan principle offers possibility to constrain error accumulated during evolution %demonstrate that during time evolution SPDM (Single Particle Density Matrix) potentially becomes singular (noninvertible), Here, critically examining Dirac-Frenkel's principle and the TDVP, and adopting alternatively McLachlan's principle for truncated many-body evolution, we improve on the previous multi-configurational Hartree methods, and solve the singularity problem of a noninvertible SPDM. In the process, we will also validate the resulting equations of MCTDHB in a different manner, however additionally offering a straightforward handling of the exceptional evolution points related to the singularity of the SPDM. ## Variational Principles Let us now discuss the possible variational principles in more detail. We are aiming at finding an approximate solution of the many-body Schr\ddot{o}dinger equation when the state |\Phi\rangle is restricted (or truncated). McLachlan's principle , which was presented in 1963 as a new version of Frenkel's principle, requires the minimization of the error or remainder of this approximate solution from the exact evolution. The time evolution of any state is dictated by Schr\ddot{o}dinger's equation, i\partial_t |\Phi\rangle = \hat{H} |\Phi\rangle. In other words, the evolution of state is determined by the Hamiltonian at any moment. But to make the state |\Phi\rangle manipulable, we are generally forced to restrict or confine the state |\Phi\rangle into some simple and computationally feasible forms. With the state |\Phi\rangle in restricted form, [i\partial_t - \hat{H}] |\Phi\rangle cannot be exactly zero in general. Therefore, McLachlan's principle aims at finding the approximate solution which minimizes the positive semidefinite error measure \langle\Phi| [i\partial_t - \hat{H}]^{\dagger}[i\partial_t - \hat{H}] |\Phi\rangle. The details of the corresponding procedure will be rephrased in section , after introducing a concrete way to restrict the many-body state in a computationally feasible form. Hence it is guaranteed that the equation of motion obtained from McLachlan's principle follows the exact evolution most similarly under given constraints. The most appealing feature of McLachlan's principle is thus that it is a quite intuitive principle. Since it offers the possibility to evaluate the error directly, we can intermediately, monitoring the error, increase the number of orbitals in the truncated field operator expansion, i.e. truncate the state less, to assure accuracy of the result. Alternatively, to save computational costs and time, the number of orbitals can also be decreased intermediately, i.e. by truncating the state more, in particular in cases where decreasing the number of orbitals does not affect the accuracy of result significantly. Our scheme, described in detail below, in which McLachlan's principle is applied, therefore offers the opportunity to dynamically adjust the truncation of field operators or the state itself properly during computational time evolution, since we can monitor the error. As a particular benefit, a convergence test, mandatory for MCTDHB, is unnecessary, as we can directly obtain an error which indicates automatically how well our approach describes the interacting system of bosons. On the other hand, the TDVP which was carried out in MCTDHB requires stationarity of action \delta S=0. This does not necessarily mean an extremum (minimum or maximum) of the action. Though stationary points include local extrema, the principle practically imposes only stationarity of the action: The equation of motion comes from stationary point of the action, which is not even necessarily a local minimum or maximum. #####/ %\delta S=0 does not necessarily give extremum In many-body quantum mechanics, the action is written in terms of an expectation value of an operator-valued functional: \begin{split} S = &\int \mathrm{d} t \int \mathrm{d} \vec{x} ~ \big\langle \Phi \big| \bigg[ \frac{1}{2m} \sum_{i=1}^{3} \big( \partial_i \hat{\Psi}^{\dagger} (\vec{x}) \big) \big( \partial_i \hat{\Psi} (\vec{x}) \big) + V_{\rm trap} (\vec{x},t) \hat{\Psi}^{\dagger} (\vec{x}) \hat{\Psi} (\vec{x}) \bigg] \big| \Phi \big\rangle \\ &+ \frac{1}{2} \int \mathrm{d} t \iint \mathrm{d} \vec{x}_{\alpha} \mathrm{d} \vec{x}_{\beta} ~ \big\langle \Phi \big| V (\vec{x}_{\alpha}, \vec{x}_{\beta}) \hat{\Psi}^{\dagger} (\vec{x}_{\alpha}) \hat{\Psi}^{\dagger} (\vec{x}_{\beta}) \hat{\Psi} (\vec{x}_{\beta}) \hat{\Psi} (\vec{x}_{\alpha}) \big| \Phi \big\rangle - \int \mathrm{d} t ~ \big\langle \Phi \big| \bigg[ \frac{\big[ i \partial_t \big] + \big[ i \partial_t \big]^{\dagger}} {2} \bigg] \big| \Phi \big\rangle \end{split} which is in quantum mechanical correspondence to the classical Lagrangian action. Here, V_{\rm trap}(\vec{x},t) is the (in general time-dependent) scalar trap potential confining the atoms, V (\vec{x}_{\alpha}, \vec{x}_{\beta}) is the two-body interaction potential, and m the mass of bosons. This is a real-valued functional of the many-body state | \Phi \rangle . Sometimes the action is simply expressed as S = \int \mathrm{d} t ~ \langle \Phi | \hat{H} - i \partial_t | \Phi \rangle. But this is not a proper expression. As the Lagrangian L(q,\dot{q};t) in classical mechanics is variable-sensitive, the proper form of the action in many-body quantum mechanics is important problem. Variationally changing the state | \Phi \rangle and the temporal change of the state \partial_t | \Phi \rangle, we find the evolution of the state around the stationary action point. But here a problem occurs. When the form of the state is restricted or truncated for computational reasons in the sense that the state resides in a sub-Hilbert space, it is questionable whether the equation of motion obtained in the sub-Hilbert space leads to an evolution most similarly to the exact one obtained with the unlimited full Hilbert space. Even though the equation of motion obtained variationally with a non-truncated state can give the correct many-body Schr\ddot{o}dinger equation, we cannot rely on the correctness of the equation in case the state is truncated. For example, the path of stationary action with limitations imposed on the path can in principle deviate far from the one obtained without any constraints on the path. As a simple specific example, when we restrict the state to have only an overall phase change, i.e. | \Phi (t) \rangle = e^{-i \Omega t} | \Phi \rangle, the action becomes S = \int \mathrm{d} t \big( \langle \hat{H} \rangle - \Omega \big). Depending on the value of \Omega, the action can be positive or negative. Actually there is no upper bound and lower bound on this action. Even worse is that, for some types of constrained states as above, there can be no stationary point of the action at all. So when applying the TDVP, at least the convergence upon increasing the number of orbitals, i.e. loosening the constraints, must be tested for every specific problem, to ensure the validity of the results. This is because, in contrast with McLachlan's principle, there is no direct error indicator in the TDVP which controls the accuracy of the approximation. The earliest variational principle for the approximate solution of many-body dynamics is Dirac-Frenkel's principle , which requires \langle\delta\Phi| \hat{H}-i\partial_t |\Phi\rangle = 0 , where \delta \Phi denotes possible variations of the many-body state \Phi with respect to the variational parameters. The equation is quite similar to the TDVP when the action is given by S = \int \mathrm{d}t \langle\Phi | \hat{H} - i \partial_t |\Phi\rangle. The difference and (possible) equivalences between Dirac-Frenkel's, McLachlan's and other variational principles have been extensively discussed in the past . In , it is concluded that if the relevant manifold, i. e. the sub-Hilbert space, can be parametrized by pairs of complementary parameters, the above mentioned principles are equivalent. In , it is insisted that both Dirac-Frenkel's and McLachlan's variational principles are equivalent if both \delta \Phi and \delta \Phi^* are possible independent variations. But the ``equivalence'' merely %seems to be conditional and therefore implies the same resulting equation under some given conditions, not the equivalence of the principles themselves. In addition, note that the simple-minded point of view that \delta \Phi and \delta \Phi^* are possible independent variations can easily lead to incorrect conclusions, since \delta \Phi^* is simply the complex conjugate of \delta \Phi . And therefore \delta \Phi^* is just dependent on \delta \Phi. Furthermore, as explained in detail later, principles which result in the problem of a noninvertible SPDM, which was mentioned already in the above, lack some information in comparison to McLachlan's principle, which resolves this problem. These three principles seem to be different in many senses. First of all, \delta \Phi and \delta \Phi^* are quite dependent as they are simply complex conjugate of each other. #####/ % Broe, CPL 149, 547 (1988): We show that if the manifold can be parameterized by pairs of “complementary” parameters, the abovementioned principles are equivalent. The condition of complementarity, which is a sufficient one, is demonstrated to be satisfied in a number of important applications. However, a case of non-equivalence in the recent literature warns against liberal assumptions about the equivalence of the time-dependent variational principles. #####/ % Without any confinement in a state, both give the correct time-dependent Schr\ddot{o}dinger equation of many-body state. But we should check carefully whether, whatever restrictions and conditions come to a state, both give the same equations and equivalent results. Here we are going to claim that they are not equivalent when the reduced SPDM \langle \hat{a}_{ij}^{\dagger} \rangle becomes noninvertible. \\ In summary, comparing the three variational principles, McLachlan's principle appears to be most suitable for finding a truncated many-body dynamics which approximates the real dynamics of interacting bosons. %and is thus automatically rigorously controlled. Adopting McLachlan's principle, we insist, in the following sections, that the variationally optimal truncation of the many-body dynamics of interacting bosons can be adaptively controlled with a monitored error. #####/ %Dirac-Frenkel's principle seems to have little connection to other physical principles and phenomena such as Hamilton's principle in classical mechanics or single particle quantum mechanics. %But it seems to lack a confident reasoning for the above equation. For what end should the Dirac-Frenkel principle be valid? ### Truncating a many-body state The limited or restricted forms of the state |\Phi\rangle for the truncated many-body dynamics can in principle take any form. In the simplest case, assuming that the occupation numbers of bosons concentrate in one orbital for the whole time of evolution, we can treat the many-body state with one single-particle orbital. More generally, the state will reside in a sub-Hilbert space of a specific form. An easily extendable and flexible form of the limitation on the size of the Hilbert space is the multiconfigurational time-dependent Hartree wavefunction ansatz, in which the many-body state of bosons is described as a linear combination of permanents |\vec{n}\rangle, with a finite number M of orthonormalized time-dependent single-particle orbitals. Increasing the number M of orbitals, we can easily extend the time-dependent sub-Hilbert space \mathcal{M}(t). The basic steps to be followed in the procedure are as follows. #####/ % As a result of introduction of time-varying orbitals or truncation, inverse of truncated SPDM comes in the evolution of orbitals. % As mentioned before, this hampers dynamical switching between various M (the number of orbitals) values in Dirac-Frenkel's variational approach. Firstly, the many-body Hamiltonian %which specifies the energy and time evolution of the many-body state is given by \begin{split} &\hat{H} = \int \mathrm{d} \vec{x} ~~ \hat{\Psi}^{\dagger} (\vec{x}) \bigg[ - \frac{\nabla^2} {2m} + V_{\rm trap}(\vec{x}) \bigg] \hat{\Psi} (\vec{x}) \\ &+ \frac{1}{2} \iint \mathrm{d} \vec{x}_{\alpha} \mathrm{d} \vec{x}_{\beta}   \hat{\Psi}^{\dagger} (\vec{x}_{\alpha}) \hat{\Psi}^{\dagger} (\vec{x}_{\beta}) V(\vec{x}_{\alpha}, \vec{x}_{\beta}) \hat{\Psi} (\vec{x}_{\beta}) \hat{\Psi} (\vec{x}_{\alpha}). \end{split} With a complete set of basis orbitals, the field operators of creation and annihilation of particles are expressed by the expansions \begin{split} \hat{\Psi}^{\dagger}(\vec{x}) = \sum_{i=1}^{\infty}\hat{a}_i^{\dagger} \phi_i^* (\vec{x}) ~ \quad\textrm{and}\quad \hat{\Psi}(\vec{x},t) = \sum_{i=1}^{\infty}\hat{a}_i \phi_i (\vec{x},t) . \end{split} Using these, the Hamiltonian can be written as \hat{H} = \sum_{i,j=1}^{\infty}{ \epsilon_{ij} \hat{a}_{i}^{\dagger} \hat{a}_{j} } + \frac{1}{2} \sum_{i,j,k,l=1}^{\infty}{ V_{ijkl} \hat{a}_{i}^{\dagger} \hat{a}_{j}^{\dagger} \hat{a}_{k} \hat{a}_{l} } where the single-particle matrix elements are \epsilon_{ij} = \int \mathrm{d} \vec{x} ~~ \phi_i^* (\vec{x}) \bigg[ -\frac{\nabla^2}{2m} + V_{\rm trap} (\vec{x}) \bigg] \phi_j (\vec{x}) , while the two-body interaction elements are represented by V_{ijkl} = \iint \mathrm{d} \vec{x}_{\alpha} \mathrm{d} \vec{x}_{\beta} ~ \phi_i^* (\vec{x}_\alpha) \phi_j^* (\vec{x}_\beta) V(\vec{x}_\alpha, \vec{x}_\beta) \phi_k (\vec{x}_\beta) \phi_l (\vec{x}_\alpha) . We abbreviate sometimes, for the sake of convenience, \begin{split} &\hat{a}_{ij}^{\dagger} \equiv \hat{a}_{i}^{\dagger} \hat{a}_{j}, \quad %% \hat{a}_{ijkl}^{\dagger\dagger} \equiv \hat{a}_{i}^{\dagger} \hat{a}_{j}^{\dagger} \hat{a}_{k} \hat{a}_{l}, \\ &\hat{a}_{ijk}^{\dagger\dagger} \equiv \hat{a}_{i}^{\dagger} \hat{a}_{j}^{\dagger} \hat{a}_{k}, \quad %% \hat{a}_{ijk}^{\dagger} \equiv \hat{a}_{i}^{\dagger} \hat{a}_{j} \hat{a}_{k} \end{split} and so on. Then, the Hamiltonian can be written in short-hand form as \hat{H} = \hat{h} + \frac{1}{2}\hat{V} = \sum_{i,j=1}^{\infty} \epsilon_{ij} \hat{a}_{ij}^{\dagger} + \frac{1}{2} \sum_{i,j,k,l=1}^{\infty} V_{ijkl} \hat{a}_{ijkl}^{\dagger\dagger}. To find the many-body ground state |G\rangle, we have to minimize the energy expectation value E_G = \langle G|\hat{H}|G\rangle. Without any restrictions on the state |G\rangle, the actual exact ground state will be found. However realistically, we cannot describe a state exactly as this will in general require infinitely many orbitals (or variables). So we confine the state into a space of finite dimension, i.e. using only finite number M of orbitals, which is computationally feasible, | G^{\mathcal{M}} \rangle = \sum_{\vec{n} \in \mathcal{M}} C_{\vec{n}} | \vec{n} \rangle. The above is regarded as the truncation of the many-body bosonic state. Here |\vec{n} \in \mathcal{M}\rangle indicates a normalized Fock state or a permanent |\vec{n} \rangle \equiv \frac{ \big(\hat{a}_1^{\dagger}\big)^{n_1} \big(\hat{a}_2^{\dagger}\big)^{n_2} \cdots \big(\hat{a}_M^{\dagger}\big)^{n_M}} {\sqrt{n_1! n_2! \cdots n_M!}} |\textrm{vac} \rangle \quad\text{with}\quad \sum_{i=1}^{M} n_i = N, which is a N particle state of which the individual members are composed of n_1 particles in \phi_1 (\vec{x}) orbital, n_2 particles in \phi_2 (\vec{x}) orbital, \cdots, and n_M particles in \phi_M (\vec{x}) orbital. These M orbitals must be orthonormalized to each other. \int \mathrm{d} \vec{x} ~ \phi_i^* (\vec{x}) \phi_j (\vec{x}) = \delta_{ij}. These orbitals compose the sub-Hilbert manifold spanned by M orthonormal orbitals, which will be denoted by \mathcal{M} in our context. The operators of creation and annihilation on these are related to the field operators in position space by the inversion of Eq. , \hat{a}_i^{\dagger} = \int \mathrm{d} \vec{x} ~ \hat{\Psi}^{\dagger} (\vec{x}) \phi_i (\vec{x}) \quad\textrm{and}\quad %% \hat{a}_i = \int \mathrm{d} \vec{x} ~ \hat{\Psi} (\vec{x})\phi_i^* (\vec{x}) . These primitive steps are well explained in . See those for a detailed account. Here, a sub-Hilbert space spanned by \sum_{i=1}^{M} c_{i} \phi_{i} (\vec{x}) is to be chosen so as to describe the ground state optimally. And the coefficients C_{\vec{n}} which give minimum of energy are to be determined, too. Then, |G^{\mathcal{M}}\rangle can be considered as optimal truncation of the actual ground state |G\rangle. The details on how to proceed concretely will follow below in section . In the time-evolving case, we change not only the coefficients C_{\vec{n}} along time, but also the sub-Hilbert space \mathcal{M} changes with time. With a time-varying truncation of the many-body state, we can express the state as | \Phi (t) \rangle = \sum_{\vec{n} \in \mathcal{M}(t)} C_{\vec{n}} (t) | \vec{n} ; t \rangle , with time-varying orbitals and their conjugate creation operators \hat{a}_i^{\dagger} (t) = \int \mathrm{d} \vec{x}~ \hat{\Psi}^{\dagger} (\vec{x}) \phi_i (\vec{x}, t) . This approach was also incorporated in MCTDHB . The time differentiation of the state | \Phi (t) \rangle then becomes \begin{split} i \partial_t | \Phi (t) \rangle &= \sum_{\vec{n} \in \mathcal{M}(t)} \Big[ \big( i \partial_t C_{\vec{n}} (t) \big) | \vec{n} ; t \rangle \\ &+ C_{\vec{n}} (t) \sum_{i=1}^{M} \int \mathrm{d} \vec{x} ~ \big( i \partial_t \phi_i (\vec{x}, t) \big) \hat{\Psi}^{\dagger} (\vec{x}) \hat{a}_{i} | \vec{n} ; t \rangle \Big]. \end{split} Expanding i \partial_t \phi_i (\vec{x}, t) with a complete basis yields i \partial_t \phi_i (\vec{x}, t) = \sum_{k=1}^{\infty} \phi_{k} (\vec{x}, t) ~ t_{ki} (t) . where the matrix t_{ki} for 1 \leq k \leq M indicates an {\em inner rotation} inside the sub-Hilbert space, whereas t_{ki} for k > M changes sub-Hilbert space itself. Integrating both sides after multiplying with \phi_{k}^{*} (\vec{x},t), we obtain t_{ki} (t) = \int \mathrm{d}\vec{x} ~ \phi_{k}^{*} (\vec{x},t) \big( i \partial_t \phi_i (\vec{x},t) \big). Then Eq.  can be expressed in the alternative form \begin{split} i \partial_t | \Phi (t) \rangle = \sum_{\vec{n} \in \mathcal{M}(t)} \Big[ &\big( i \partial_t C_{\vec{n}} (t) \big) | \vec{n} ; t \rangle \\ &+ C_{\vec{n}} (t) \sum_{k=1}^{\infty} \sum_{i=1}^{M} t_{ki} \hat{a}_{ki}^{\dagger} | \vec{n} ; t \rangle \Big]. \end{split} These preliminaries will be used in the following sections, by providing the tools for describing the truncated state |\Phi\rangle optimally. #####/ %% Since we do not count orbitals outside M, \phi_{k}'s for k > M are just pictorial. The actual calculation takes upto M. %Thus show nonequivalence of Dirac-Frenkel and MacLachlan %Uniqueness of the solution ###[#subsec-Adapting-the-number-of-orbitals] Adapting the number of orbitals The governing equation of MCTDHB which comes from either Dirac-Frenkel's principle or TDVP implies that the SPDM must be always invertible, not only initially but also at any instant time afterwards. We quote here the equation (26) in , which is \begin{split} &i \partial_t \phi_{k} (\vec{x},t) \\ &= \sum_{l=M+1}^{\infty} \phi_{l} (\vec{x},t) \bigg[ \epsilon_{lk} + \sum_{i,n,p,q=1}^{M} V_{lnpq} \langle \rho \rangle^{-1}_{ki} \langle \hat{a}_{inpq}^{\dagger\dagger} \rangle \bigg] \end{split} in our notation, where \langle \rho \rangle^{-1}_{ki} represents the inverse of the SPDM \langle \rho_{ij} \rangle \equiv \langle \hat{a}_{ij}^{\dagger} \rangle. In , it seems to be taken for granted that the SPDM is always invertible. When the SPDM becomes noninvertible, however, the MCTDHB method in fact suddenly fails. %Isn't this inevitable? Furthermore, why in the first place, with MCTDHB, should we be unable to propagate pure %single condensate into fragmented state? What was wrong? or what was missing? #####/ %When the reduced SPDM \langle \hat{a}_{ij}^{\dagger} \rangle becomes noninvertible, MCTDHB which comes from %Dirac-Frenkel's principle fails. For example, when the whole bosons of our interest reside initially in single orbital, MCTDHB cannot propagate this pure condensate into any fragmented state. The method does not provide a way to find the form of the second orbital in this case. To decrease the sub-Hilbert space or the number of orbitals is not an issue. We can simply eliminate those orbitals with an occupation number ignorable, i.e.  not of order N. However, to increase the sub-Hilbert space dimension, that is the number of orbitals M, becomes difficult since an increase of the sub-Hilbert space and an additional orbital must be set up optimally. Noninvertibility of the SPDM can obviously also happen dynamically. Unoccupied orbitals or scarcely occupied orbitals thus cause a problem with the dynamics of MCTDHB. These problems cannot be resolved by Dirac-Frenkel's principle or the TDVP. As McLachlan's principle is based on requiring that the least error should be acquired during time evolution, we may resolve this problem by finding an additional orbital which minimizes the error, as will be expounded in detail below. #####/ %Applying McLachlan's alternatively, we have tried to solve these problems and to see things in a different point of view. %noninvertibility of single-particle density matrix %during time evolution hampers dynamical switching between various M values for Dirac-Frenkel %Explain why not a problem for MacLachlan McLachlan's principle does not use the action, but the concept of error or remainder. We know the exact form of the many-body Schr\ddot{o}dinger equation. It is i \partial_t | \Phi \rangle = \hat{H} | \Phi \rangle with the full Hamiltonian Eq.  in second quantization form. Since the exact calculation is too cumbersome, we can represent the state by the relatively simple multiconfigurational time-dependent Hartree form. The remainder from exact evolution in any case becomes \big[ i \partial_t - \hat{H} \big] \big| \Phi \big\rangle . In this expression, the left part i \partial_t | \Phi \big\rangle is an evolution of the state |\Phi\rangle in its limited, truncated form and the right part \hat{H} | \Phi \big\rangle represents the exact evolution. Here, the initial state is specified in its truncated form, while the Hamiltonian \hat{H} is not truncated. A quantitative measure of the instantaneous error is then \big\langle \Phi \big| \big[ i \partial_t - \hat{H} \big]^{\dagger} \big[ i \partial_t - \hat{H} \big] \big| \Phi \big\rangle \geq 0 , which is (by definition) positive semidefinite. We have to minimize this error by varying the truncated state in Eq. . ##[#sec-Minimizing-the-energy] Minimizing the energy of the truncated many-body state ### Variational method with Lagrange multipliers Before investigating the dynamics, let us demonstrate the time-independent scheme first. To find the many-body ground state, we have to minimize the energy expectation value E_{G^{\mathcal{M}}} = \langle G^{\mathcal{M}}| \hat{H} |G^{\mathcal{M}}\rangle = \Big[ \sum_{\vec{m} \in \mathcal{M}} \langle \vec{m} | C_{\vec{m}}^* \Big] \hat{H} \Big[ \sum_{\vec{n} \in \mathcal{M}} C_{\vec{n}} | \vec{n} \rangle \Big] by varying the coefficients C_{\vec{n}}'s and the set of M orbitals \{\phi_i\}, subject to the (1+M^2) constraints \sum_{\vec{n}} C_{\vec{n}}^* C_{\vec{n}} = 1 \qquad \text{(1 constraint)} and \begin{split} \int \mathrm{d} \vec{x} ~ \phi_i^* (\vec{x}) \phi_j (\vec{x}) = \delta_{ij} \quad \textrm{for }i, j \leq M \\ \text{(M^2 constraints)}. \end{split} The variation with respect to the expansion coefficients C_{\vec{n}} gives \frac{\partial E_{G^{\mathcal{M}}}} {\partial C_{\vec{n}}^*} = \langle \vec{n} | \hat{H} \sum_{\vec{m} \in \mathcal{M}} C_{\vec{m}} | \vec{m} \rangle = \lambda C_{\vec{n}} = E_{G^{\mathcal{M}}} C_{\vec{n}}, where \vec{n} \in M . We have a real functional E_{G^M} Eq.  to be minimized, and one real equation of constraint Eq. , with \frac{(N+M-1)!}{N!(M-1)!} complex variables C_{\vec{n}} when we fix the total number of bosons at N. The undetermined Lagrange multiplier \lambda which has to be real is determined to be E_{G^{\mathcal{M}}} with the help of constraint Eq. . Using the properties \epsilon_{ij}^* = \epsilon_{ji}, V_{ijkl} = V_{jilk} and V_{ijkl}^* = V_{lkji} = V_{klij}, we can express the above equation explicitly as \begin{split} &\langle \vec{n} | \hat{H} \sum_{\vec{m} \in \mathcal{M}} C_{\vec{m}} | \vec{m} \rangle = E_{G^{\mathcal{M}}} C_{\vec{n}} \\ &= \sum_l \bigg[ \epsilon_{ll} + \frac{1}{2} V_{llll} (n_l-1) \\ &~~~~~~~~~~~~~~~~~~~ + \frac{1}{2} \sum_{k}{}^{'} (V_{lklk} + V_{lkkl}) n_k \bigg] n_l C_{\vec{n}} \\ &~+ \sum_{j,l}{}^{'} \bigg[ \epsilon_{lj} + V_{lllj} (n_l-1) + V_{ljjj} n_j \\ &~~~~~~~~~~~~~~~~~~~ + \sum_{k}{}^{'} (V_{lkkj} + V_{lkjk}) n_k \bigg] \sqrt{(n_j+1) n_l} C_{\vec{n}^{j}_{l}} \\ &~~+ \frac{1}{2} \sum_{j,l}{}^{'} V_{lljj} \sqrt{(n_j+2)(n_j+1)(n_l-1)n_l} C_{\vec{n}^{jj}_{ll}} \\ &~~+ \frac{1}{2} \sum_{i,j,l}{}^{'} V_{llji} \sqrt{(n_i+1)(n_j+1)(n_l-1)n_l} C_{\vec{n}^{ij}_{ll}} \\ &~~+ \frac{1}{2} \sum_{j,k,l}{}^{'} V_{lkjj} \sqrt{(n_j+2)(n_j+1)n_k n_l} C_{\vec{n}^{jj}_{kl}} \\&~~ + \frac{1}{2} \sum_{i,j,k,l}{}^{'} V_{lkji} \sqrt{(n_i+1)(n_j+1)n_k n_l} C_{\vec{n}^{ij}_{kl}}, \end{split} where the primed summation \sum' is performed such that different indices can have only different values. For example, summations over two and three indices are \sum_{j,l}{}^{'} \equiv \sum_{j} \sum_{l \neq j}, \sum_{i,j,l}{}^{'} \equiv \sum_{i} \sum_{j \neq i} \sum_{l \neq (i,j)}, and \sum_{k}{}^{'} inside a bracket means \sum_{k \neq \text{(the other indices)}}. Though Eq.  is just an eigenvalue equation with fixed matrix components, the terms \epsilon_{ij} and V_{ijkl} are matrix elements depending on the orbitals, which are also to be tuned as follows. For variation with respect to the orbitals \{\phi_k\}, functional differentiation is used. As the permanent | \vec{n} \rangle is constructed from repeatedly applying creation operators of particles in the M orbitals, it can be regarded to be given by multiple integrations over the orbitals \phi_k (\vec{x}), |\vec{n}\rangle = \int \mathrm{d}\vec{x}_{\alpha} \phi_{i} (\vec{x}_{\alpha}) \hat{\Psi}^{\dagger} (\vec{x}_{\alpha}) \int \mathrm{d}\vec{x}_{\beta} \phi_{j} (\vec{x}_{\beta}) \hat{\Psi}^{\dagger} (\vec{x}_{\beta}) \cdots |\textrm{vac}\rangle . As each functional differentiation contributes to the result, this will be counted by a factor \hat{a}^{\dagger}_{k}, which results in \frac{\partial \langle \Phi |} {\partial \phi_k^* (\vec{x})} = \langle \Phi | \hat{a}_{k}^{\dagger} \hat{\Psi} (\vec{x}). But as the full hamiltonian \hat{H}, which is given in field form by Eq. , is independent on the M orbitals chosen, the functional differentiation of \hat{H} with respect to the orbitals \{\phi_k\} gives zero, \frac{\partial \hat{H}} {\partial \phi_k^* (\vec{x})} = 0. Then functional variation of E_{G^{\mathcal{M}}} Eq.  with respect to the orbitals \{\phi_k\}, combined with the functional constraints Eq. , leads to \begin{split} &\frac{\partial E_{G^{\mathcal{M}}}}{\partial \phi_k^* (\vec{x})} = \\ &\Big[ \sum_{\vec{m} \in \mathcal{M}} \langle \vec{m} | C_{\vec{m}}^* \Big] \hat{a}_k^{\dagger} \hat{\Psi} (\vec{x}) \hat{H} \Big[ \sum_{\vec{n} \in \mathcal{M}} C_{\vec{n}} | \vec{n} \rangle \Big] = \sum_{j=1}^{M} \lambda_{kj} \phi_j (\vec{x}) , \end{split} where \lambda_{jk} = \lambda_{kj}^* is Hermitian matrix. Here the method of Lagrange multipliers with complex functional variables is used . Integrating each side over space after multiplication with \phi_{l}^* (\vec{x}) yields \Big[ \sum_{\vec{m} \in \mathcal{M}} \langle \vec{m} | C_{\vec{m}}^* \Big] \hat{a}_k^{\dagger} \hat{a}_l \hat{H} \Big[ \sum_{\vec{n} \in \mathcal{M}} C_{\vec{n}} | \vec{n} \rangle \Big] = \begin{cases} \lambda_{kl} & \textrm{for } l \leq M \\ 0 & \textrm{for } l > M . \end{cases} Here, using Eq.  and the property that \vec{m}_{k}^{l} \in \mathcal{M} when k,l \leq M and \vec{m} \in \mathcal{M} , the undetermined set of Lagrange multipliers \lambda_{kl} becomes related to E_{G^{\mathcal{M}}} by \lambda_{kl} = E_{G^{\mathcal{M}}} \langle \hat{a}_{kl}^{\dagger} \rangle for k, l \leq M. Using Eq.  and the bosonic commutation relations between field operators, i.e. [\hat{\Psi}(\vec{x}), \hat{\Psi}^{\dagger}(\vec{x}')] = \delta \big( \vec{x} - \vec{x}' \big), [\hat{\Psi}(\vec{x}), \hat{a}_{k}^{\dagger}] = \phi_k (\vec{x}), and so on, Eq.  \langle \hat{a}_k^{\dagger} \hat{\Psi} (\vec{x}) \hat{H} \rangle = \sum_{j=1}^{M} \lambda_{kj} \phi_j (\vec{x}) is explicitly expressed as \begin{split} &\sum_{j=1}^{\infty} \langle \hat{a}_{kj}^{\dagger} \rangle \bigg[ - \frac{ \nabla^2} {2m} + V_{\rm trap}(\vec{x}) \bigg] \phi_j (\vec{x}) \\ &+ \sum_{p,q,j=1}^{\infty} \langle \hat{a}_{kpqj}^{\dagger\dagger} \rangle \int \mathrm{d} \vec{x}' ~ \phi_p^* (\vec{x}') V(\vec{x}, \vec{x}') \phi_q (\vec{x}') \phi_j (\vec{x}) \\ &+ \langle \hat{a}_k^{\dagger} \hat{H} \hat{\Psi} (\vec{x}) \rangle = \sum_{j=1}^{M} \lambda_{kj} \phi_j (\vec{x}) . \end{split} Since we can eliminate the annihilations above M, we obtain \begin{split} &\sum_{j=1}^{M} \langle \hat{a}_{kj}^{\dagger} \rangle \hat{h} \phi_j (\vec{x}) + \sum_{p,q,j=1}^{M} \langle \hat{a}_{kpqj}^{\dagger\dagger} \rangle \hat{V}_{pq} \phi_j (\vec{x}) \\ &= \sum_{j=1}^{M} \Big[ \lambda_{kj} - \langle \hat{a}_k^{\dagger} \hat{H} \hat{a}_j \rangle \Big] \phi_j (\vec{x}) \\ &= \sum_{j=1}^{M} \Big[ E_{G^M} \langle \hat{a}_k^{\dagger} \hat{a}_j \rangle - \langle \hat{a}_k^{\dagger} \hat{H} \hat{a}_j \rangle \Big] \phi_j (\vec{x}) \equiv \sum_{j=1}^{M} \tilde{\lambda}_{kj} \phi_j (\vec{x}), \end{split} where the new undetermined Lagrange multipliers \tilde{\lambda}_{kj} \equiv \lambda_{kj} - \langle \hat{a}_k^{\dagger} \hat{H} \hat{a}_j \rangle which satisfy \tilde{\lambda}_{kj}^* = \tilde{\lambda}_{jk} are introduced. We abbreviated, for the sake of convenience, two single-particle and interaction operators defined by \hat{h} \phi_j (\vec{x}) \equiv \bigg[ - \frac{\nabla^2} {2m} + V_{\rm trap}(\vec{x}) \bigg] \phi_j (\vec{x}) and \hat{V}_{pq} \phi_j (\vec{x}) \equiv \int \mathrm{d} \vec{x}' ~ \phi_p^* (\vec{x}') V (\vec{x}, \vec{x}') \phi_q (\vec{x}') \phi_j (\vec{x}) . ### Method of steepest constrained descent To find the tentative ground state | G^{M} \rangle, we have to find the coefficient C_{\vec{n}}'s and the complex orbital function \phi_k (\vec{x})'s satisfying Eq.  and simultaneously. Additionally, the solutions must satisfy all (1 + M^2) constraints Eq. , . The number of real values which we should find is 2\frac{(N+M-1)!}{N!(M-1)!} for the set of the C_{\vec{n}}, 2M real functions for the \phi_k (\vec{x}), and (1 + M^2) real values for the undetermined Lagrange multiplier E_{G^M} and the \lambda_{kj}. The number of given real equations therefore is 2\frac{(N+M-1)!}{N!(M-1)!} for Eq. , 2M real functional equations for Eq. , and (1 + M^2) for Eq. , . Apart from the large number of variables, what makes matters even more complicated is that the equations , , ,  are coupled together. To find a self-consistent solution is therefore obviously a very difficult problem. #####/ %is much more difficult problem than finding equations that the solution must satisfy. %So we have to know, in addition, the way to find |G^M \rangle, actually C_{\vec{n}}'s and \phi_k (\vec{x})'s from given equations. In ref. , the authors started from an initial guess, then iteratively, with a convergence check, they obtained a solution. A few years later in ref. , applying the Wick rotation it \rightarrow \tau on the equations of motion, they introduced the so-called imaginary time propagation. They claimed that this reduces any arbitrary initial many-body state after a sufficient time of propagation to the ground state. The imaginary time evolution i \frac{\partial}{\partial t} |\Phi \rangle = \hat{H} |\Phi \rangle \Rightarrow - \frac{\partial}{\partial \tau} |\Phi \rangle = \hat{H} |\Phi \rangle implies that e^{- i \hat{H} t } \Rightarrow e^{- \hat{H} \tau } |\Phi \rangle. As \tau goes to infinity, this seems to indicate that only the ground state survives and the excited states would no longer contribute. The contribution of the excited states decays exponentially according to a factor which is proportional to energy difference from the ground state and \tau. #####/ %(The reason why we introduce another method.) %No flexibility. %Orbitals must be orthonormalized. But I think imageinary time propagation does not guarantee it. Although normalization of C_n can change, keeping orbitals' orthonormality might be safe. Here, we present the method of steepest constrained descent, of which the understanding is fundamental to the theory and design of higher-order algorithms of optimization with constraints. Since our problem requires optimization with constraints on the variables Eq. , , a simple-minded gradient (or, as it is more commonly termed steepest) descent method does not work in our case. There are already numerous higher-order optimization algorithms without constraints such as Newton's method, conjugate gradient method, BFGS (Broyden-Fletcher-Goldfarb-Shanno) method, and the Barzilai-Borwein method . But these methods are not applicable to our problem since the function to be minimized with constraints on the variables is generally unbounded without constraints. For example, the energy \langle \hat{H} \rangle with unnormalized state | \Phi \rangle can be zero, or \pm\infty. So when the actual ground-state energy is positive, the global minimum is where all C_{\vec{n}} = 0. For this case, e.g. Newton's method will simply lead to an incorrect solution. In other words, since our final point is not a global minimum without constraints, it can not be approximated as a quadratic function \big[ f(\bf{x}) = \frac{1}{2} \bf{x}^{T} A \bf{x} + \bf{b}^{T} \bf{x} \big] around the minimum point (with constraints included). So all of these above-mentioned other methods, which are based on the expansion properties of the function around the minimum to second order will fail. #####/ %Anyway the understanding of steepest constrained descent is fundamental to the theory and design of higher-order algorithms of optimization with constraints. On the other hand, the method of steepest constrained descent guarantees that any given state is propagated to the neighboring lowest value point along the steepest constrained path for given variational parameters. Though it propagates a state only to a local neighboring minimum point, we can find in many cases the global minimum, or the ground state, from a well-chosen initial state, and with in addition well-chosen variational parameters. Although this does not deliver the state along the shortest path, the very large number of degrees of freedom on the choice of variational parameters or the sequential processing (separation) of variations can compensate it in many cases. Furthermore the step size can be determined by one-point or two-point method. The flexibility of the method of steepest constrained descent will therefore be beneficial for finding the ground state. Let us see the process in more detail. The first step is to find (by educated guess) an appropriate initial state, specifying the coefficients C_{\vec{n}} and the M orbitals which we believe are appropriate to approximately describe the ground state of a given system. %It doesn't have to be accurate, though good guess would be definitely better to save time and calculation costs. From this initial guess, the state is propagated as follows. For the expansion coefficient C_{\vec{n}}'s, using the steepest constrained descent , \frac{d C_{\vec{n}}} {d \tau} = - \Delta_{C}(\tau) \bigg[ \langle \vec{n} | \hat{H} \sum_{\vec{m}} C_{\vec{m}} | \vec{m} \rangle - \lambda C_{\vec{n}}, \bigg] where \Delta_{C}(\tau) is any arbitrary positive function of \tau which is introduced to satisfy \sum_{\vec{n}} \frac{d C_{\vec{n}}^*}{d\tau} \frac{d C_{\vec{n}}}{d\tau}= const at a certain given instant \tau and therefore can be chosen in a convenient way to save time and calculation costs. Since it must satisfy \sum_{\vec{n}} C_{\vec{n}}^* C_{\vec{n}} = 1, i.e. \sum_{\vec{n}} \big[ C_{\vec{n}}^* \frac{d C_{\vec{n}}} {d \tau} + \frac{d C_{\vec{n}}^*} {d \tau} C_{\vec{n}} \big] = 0, \lambda is to be \langle \hat{H} \rangle. As another option, we can use a polar representation for C_{\vec{n}}. Representing the complex variable C_{\vec{n}} by an Euler representation with a radius \xi_{\vec{n}} and an angle \theta_{\vec{n}} gives C_{\vec{n}} = \xi_{\vec{n}} e^{i \theta_{\vec{n}}}. Then the constraint Eq.  becomes \sum_{\vec{n}} \xi_{\vec{n}}^2 = 1, restricting only the radial component of the complex variables C_{\vec{n}}. Since we do not have to confine the variable change into the specific form \sum_{\vec{n}} \big[ \big( \frac{d \xi_{\vec{n}}}{d\tau} \big)^2 + \xi_{\vec{n}}^2 \big( \frac{d \theta_{\vec{n}}}{d\tau} \big)^2 \big] = const, separating the two variable sets can be much more efficient in this case. That is, we set \frac{d \xi_{\vec{n}}}{d\tau} = 0 and \sum_{\vec{n}} \big( \frac{d \theta_{\vec{n}}}{d\tau} \big)^2 = const for some given \tau time spans which would be chosen in computationally favorable way, and \frac{d \theta_{\vec{n}}}{d\tau} = 0 and \sum_{\vec{n}} \big( \frac{d \xi_{\vec{n}}}{d\tau} \big)^2 = const for the other \tau spans. Then the method of steepest constrained descent gives \frac{d \xi_{\vec{n}}} {d \tau} = 0 , \qquad %% \frac{d \theta_{\vec{n}}} {d \tau} = - \Delta_{\theta}(\tau) \Im \big( \xi_{\vec{n}} e^{- i \theta_{\vec{n}}} \langle \vec{n} | \hat{H} \rangle \big), for some given \tau spans, and \frac{d \theta_{\vec{n}}} {d \tau} = 0 , \qquad %% \frac{d \xi_{\vec{n}}} {d \tau} = - \Delta_{\xi}(\tau) \Big[ \Re \big( e^{- i \theta_{\vec{n}}} \langle \vec{n} | \hat{H} \rangle \big) - \lambda \xi_{\vec{n}} \Big] for the other following \tau-intervals. Here \Re means real part of complex number and \Im means imaginary part of complex number. Using \sum_{\vec{n}} \xi_{\vec{n}}^2 = 1, and therefore \sum_{\vec{n}} 2 \xi_{\vec{n}} \frac{d \xi_{\vec{n}}} {d \tau} = 0, then \lambda becomes \lambda = \langle \hat{H} \rangle too. Dealing with \xi_{\vec{n}} and \theta_{\vec{n}} separately, we propagate the two variable sets successively and iteratively until convergence is achieved. With any sequence, they will finally approach the minimum. %With good choice of variables and topology, it will approach the minimum more quickly. \\ For the orbitals \phi_k (\vec{x}), the method of steepest constrained descent gives \begin{split} &\frac{d \phi_{k} (\vec{x})} {d \tau} = -\Delta_{\phi_k} (\tau) \Big[ \langle \hat{a}_{k}^{\dagger} \hat{\Psi} (\vec{x}) \hat{H} \rangle - \sum_{j=1}^{M} \lambda_{kj} \phi_{j} (\vec{x}) \Big] \\ &= - \Delta_{\phi_k} (\tau) \bigg[ \sum_{j=1}^{M} \langle \hat{a}_{kj}^{\dagger} \rangle \hat{h} \phi_j (\vec{x}) + \sum_{p,q,j=1}^{M} \langle \hat{a}_{kpqj}^{\dagger\dagger} \rangle \hat{V}_{pq} \phi_j (\vec{x}) \\ &~~~~~~~~~~~~~~~~~~ - \sum_{j=1}^{M} \tilde{\lambda}_{kj} \phi_j (\vec{x}) \bigg]. \end{split} If we propagate the orbitals separately one after another, % iteratively, i.e. only the k-th orbital changes within a certain period, the constraint becomes \int \mathrm{d}\vec{x}~ \phi_l^* (\vec{x}) \frac{\phi_k (\vec{x})}{d\tau} = 0 for l \neq k and \int \mathrm{d}\vec{x}~ \big( \phi_k^* (\vec{x}) \frac{\phi_k (\vec{x})}{d\tau} + \frac{\phi_k^* (\vec{x})}{d\tau} \phi_k (\vec{x}) \big) = 0. Then the undetermined Lagrange multipliers becomes \tilde{\lambda}_{kl} = \sum_{j=1}^{M} \langle \hat{a}_{kj}^{\dagger} \rangle \epsilon_{lj} + \sum_{p,q,j=1}^{M} \langle \hat{a}_{kpqj}^{\dagger\dagger} \rangle V_{lpqj} and \tilde{\lambda}_{kk} = \Re \Big( \sum_{j=1}^{M} \langle \hat{a}_{kj}^{\dagger} \rangle \epsilon_{kj} + \sum_{p,q,j=1}^{M} \langle \hat{a}_{kpqj}^{\dagger\dagger} \rangle V_{kpqj} \Big) . Separating the propagation of the orbitals, i.e. propagating the orbitals one after another independently, makes the process much easier and more comprehensible. %And the computational error which comes from finite step size can be handled by cutting the non-orthogonal part of the only propagating orbital. In a numerical implementation, the method of steepest constrained descent is performed in discrete steps. Then a projection onto the set of constraints must be employed at every discrete step. Using just a one-point step size method, we can use the Lagrange multipliers \lambda_{kj} as a degree of freedom to save computational cost, since we do not need to consider the propagation outside of constraints. As an example, we can set \lambda_{kj} =\langle \hat{a}_k^{\dagger} \hat{H} \hat{a}_j \rangle in Eq.  so that we do not need to calculate the quantities \langle \hat{a}_k^{\dagger} \hat{H} \hat{a}_j \rangle. However, to ensure fast convergence, the optimal step size is determined with a two-point or even four-point method. The role of Lagrange multipliers is crucial, then, because we mix two gradients at different points. Because of the finite step size, the propagation outside of the constraints at one point can be the direction of the steepest constrained descent at another point. The task in an implementation therefore is to combine the proper directions with appropriately chosen Lagrange multipliers at different points, to optimize the speed of convergence. ## Control of truncated many-body evolution ### Evaluating the error of truncated many-body evolution An instantaneous error is expressed by Eq. . Minimizing this error with a state change under the truncation Eq.  gives us the appropriate truncated many-body evolution. This offers, as a major benefit of the present approach, a quantitative value, the error, which indicates how accurately the truncated evolution describes the exact one. Explicitly expressing the error, we have \begin{split} &\langle \Phi | \big[ \hat{H} - i \partial_t \big]^{\dagger} \big[ \hat{H} - i \partial_t \big] | \Phi \rangle \\ &= \sum_{\vec{n} \in \mathcal{M}(t)} \bigg[ \langle \vec{n} | C_{\vec{n}}^* \hat{H} + \langle \vec{n} | \big(i \partial_t C_{\vec{n}}^* \big) \\ &~~~~~~~~~~~~~~~~ + \sum_{i=1}^{M} \langle \vec{n} | C_{\vec{n}}^* \hat{a}_{i}^{\dagger} \int \mathrm{d} \vec{x} \big(i \partial_t \phi_i^* (\vec{x},t) \big) \hat{\Psi} (\vec{x}) \bigg] \\ &~ \times \sum_{\vec{m} \in \mathcal{M}(t)} \bigg[ \hat{H} C_{\vec{m}} | \vec{m} \rangle - \big(i \partial_t C_{\vec{m}} \big) | \vec{m} \rangle \\ &~~~~~~~~~~~~~~~~ - \sum_{j=1}^{M} \int \mathrm{d} \vec{x}' \big(i \partial_t \phi_j (\vec{x}',t) \big) \hat{\Psi}^{\dagger} (\vec{x}') \hat{a}_{j} C_{\vec{m}} | \vec{m} \rangle \bigg] . \end{split} We minimize this instantaneous error, varying the complex variables \big( \partial_t C_{\vec{n}} \big) and \big( \partial_t \phi_i (\vec{x},t) \big) subject to the (1 + M^2) constraints \begin{split} &\partial_t \bigg[ \sum_{\vec{n}} C_{\vec{n}}^* (t) C_{\vec{n}} (t) = 1 \bigg] \\ &\Rightarrow \quad \sum_{\vec{n}} \bigg[ \big( \partial_t C_{\vec{n}}^* (t) \big) C_{\vec{n}} (t) + C_{\vec{n}}^* (t) \big( \partial_t C_{\vec{n}} (t) \big) \bigg] = 0 \end{split} and from the orthonormality condition \begin{split} &\partial_t \bigg[ \int \mathrm{d} \vec{x} \phi_i^* (\vec{x},t) \phi_j (\vec{x},t) = \delta_{ij} \bigg] \Rightarrow \quad \int \mathrm{d} \vec{x} \bigg[ \big( \partial_t \phi_i^* (\vec{x},t) \big) \\ &\times \phi_j (\vec{x},t) + \phi_i^* (\vec{x},t) \big( \partial_t \phi_j (\vec{x},t) \big) \bigg] = 0 . \end{split} The Eq.  requires probability conservation of the state itself and Eq.  requires conservation of orthonormality of orbitals. Using the expression Eq. , Eq.  can be expressed as the hermiticity condition t_{ji}^* = t_{ij}. Variation with respect to \partial_t C_{\vec{n}}^* leads to \langle \vec{n} | i \big[ \hat{H} - i \partial_t \big] |\Phi\rangle = \lambda (t) C_{\vec{n}} resulting in i \partial_t C_{\vec{n}} = \langle \vec{n} | \big[ \hat{H} - \sum_{i=1}^{\infty} \sum_{j=1}^{M} t_{ij} \hat{a}_{ij}^{\dagger} \big] |\Phi\rangle + i \lambda (t) C_{\vec{n}}. As the constraint Eq.  enforces \lambda (t) = 0, % for the time evolution \langle \vec{n} | \big[ \hat{H} - i \partial_t \big] |\Phi\rangle = 0, and therefore the time evolution of the expansion coefficients takes the form \begin{split} i \partial_t C_{\vec{n}} = \langle \vec{n} | \big[ \hat{H} - \hat{t}~ \big] |\Phi\rangle \end{split} where \vec{n} \in M(t) and \hat{t} = \sum_{i,j} t_{ij} \hat{a}_{ij}^{\dagger}. Variation with respect to \partial_t \phi_k^* (\vec{x},t) gives \langle \Phi | i \hat{a}_{k}^{\dagger} \hat{\Psi} (\vec{x}) \big[ \hat{H} - i \partial_t \big] |\Phi\rangle = \sum_{l=1}^{M} \lambda_{kl} \phi_{l} (\vec{x},t) . After multiplying both sides with \phi_l^* (\vec{x},t) and integrating, one finds \langle \Phi | i \hat{a}_{k}^{\dagger} \hat{a}_{l} \big[ \hat{H} - i \partial_t \big] |\Phi\rangle = \begin{cases} \lambda_{kl} (t) &\text{ if } l \leq M \\ 0 &\text{ if } l>M . \end{cases} Since \langle\vec{n}| \hat{a}_{kl}^{\dagger} belongs to {\mathcal M}(t) whenever \vec{n} \in \mathcal{M}(t) and k,l \leq M, all \lambda_{kl} (t) here become zero too with the help of Eq. . So for any k and l, % for the time evolution \langle\Phi | \hat{a}_{kl}^{\dagger} \big[ \hat{H} - i \partial_t \big] | \Phi \rangle = 0. Expanding the above equation for l>M, \langle \hat{a}_{kl}^{\dagger} \hat{H} \rangle = \sum_{j=1}^{M} \langle \hat{a}_{kj}^{\dagger} \rangle \int \mathrm{d}\vec{x} \phi_l^* (\vec{x},t) \big( i \partial_t \phi_j (\vec{x},t) \big) . Since the SPDM can be noninvertible, we reduce the density matrix by eliminating unoccupied orbitals in which the eigenvalues of the SPDM becomes zero (relative to {\cal O}(N) in the limit N\rightarrow\infty) after diagonalizing the SPDM. In the process of diagonalizing the SPDM, the orbitals are unitarily transformed so that the essentially unoccupied orbitals can be found and eliminated. Introducing the inverse of this reduced density matrix \langle \rho_{ki} \rangle \equiv \langle \hat{a}_{ki}^{\dagger} \rangle \equiv \langle \hat{a}_{k}^{\dagger} \hat{a}_{i} \rangle %\col{Why is there a dagger here on the \hat a_{ki}??} when the {\cal O}(N) occupied orbitals exist up to the M_1th orbital (\sum_{i=1}^{M_1} \langle \rho \rangle_{ki}^{-1} \langle \rho_{ij} \rangle = \delta_{kj}) and using the completeness relation \sum_{l=1}^{\infty} \phi_{l}^* (\vec{x}',t) \phi_{l} (\vec{x},t) = \delta (\vec{x}'-\vec{x}), the evolution equation of the orbitals acquires the form \begin{split} &i \partial_t \phi_k (\vec{x},t) = \sum_{l=1}^{M} t_{lk} \phi_{l} (\vec{x},t) + \sum_{l=M+1}^{\infty} \sum_{i=1}^{M_1} \langle \rho \rangle_{ki}^{-1} \langle \hat{a}_{i}^{\dagger} \hat{a}_{l} \hat{H} \rangle \phi_{l} (\vec{x},t) \\ &= \sum_{l=1}^{M} t_{lk} \phi_{l} (\vec{x},t) + \sum_{l=M+1}^{\infty} \sum_{i=1}^{M_1} \langle \rho \rangle_{ki}^{-1} \langle \bigg[ \sum_{n=1}^{M_1} \epsilon_{ln} \hat{a}_{in}^{\dagger} + \sum_{n,p,q=1}^{M_1} V_{lnpq} \hat{a}_{inpq}^{\dagger\dagger} \bigg] \rangle \phi_{l} (\vec{x},t) \\ &= \sum_{l=1}^{M} t_{lk} \phi_{l} (\vec{x},t) + \sum_{l=M+1}^{\infty} \bigg[ \epsilon_{lk} + \sum_{i,n,p,q=1}^{M_1} V_{lnpq} \langle \rho \rangle_{ki}^{-1} \langle \hat{a}_{inpq}^{\dagger\dagger} \rangle \bigg] \phi_{l} (\vec{x},t) \end{split} for k \leq M_1. Here, we divide the evolution of the orbitals into two parts. We call the left l\leq M part of Eq.  inner rotation and the right l>M part of Eq.  rotation toward the outside the sub-Hilbert space. Since the sub-Hilbert space spanned by M orbitals, \sum_{i=1}^{M} c_{i} \phi_{i} (\vec{x},t), does not change under inner rotation, we realize that only a rotation toward the outside deforms the sub-Hilbert space. For the evolution inside the sub-Hilbert space, i.e. inner rotation, we simply use t_{ij} defined in , which can be any Hermitian matrix. Using Eq. , , , the expression for the error Eq.  can be strongly simplified: \begin{split} &\langle \Phi | \big[ \hat{H} - i \partial_t \big]^{\dagger} \big[ \hat{H} - i \partial_t \big] | \Phi \rangle = \langle \Phi | \big[ \frac{1}{2} \sum_{i,j,k,l=1}^{\infty} V_{ijkl} \hat{a}_{ijkl}^{\dagger\dagger} \big] \big[ \hat{H} - i \partial_t \big] | \Phi \rangle = \frac{1}{2} \sum_{i,j=1}^{M_1} \sum_{k,l=1}^{\infty} V_{ijkl} \langle \Phi | \hat{a}_{ijkl}^{\dagger\dagger} \big[ \hat{H} - i \partial_t \big] | \Phi \rangle \\ &= -\!\sum_{i,j,k,n,r,s,p,q=1}^{M_1} \sum_{l=M+1}^{\infty}\! V_{ijkl} V_{lspq} \langle \hat{a}_{ijkn}^{\dagger\dagger} \rangle \langle \rho \rangle_{nr}^{-1} \langle \hat{a}_{rspq}^{\dagger\dagger} \rangle +\!\sum_{i,j,k,s,p,q=1}^{M_1} \sum_{l=M+1}^{\infty}\! V_{ijkl} V_{lspq} \langle \hat{a}_{ijskpq}^{\dagger\dagger\dagger} \rangle \\ &~~~ +\frac{1}{2}\!\sum_{i,j,p,q=1}^{M_1} \sum_{k=1}^{M} \sum_{l=M+1}^{\infty}\! V_{ijkl} V_{lkpq} \langle \hat{a}_{ijpq}^{\dagger\dagger} \rangle %\\&~~~ +\frac{1}{2} \sum_{i,j,p,q=1}^{M_1} \sum_{k=1}^{\infty} \sum_{l=M+1}^{\infty}\! V_{ijkl} V_{lkpq} \langle \hat{a}_{ijpq}^{\dagger\dagger} \rangle . \end{split} The above equation represents our main result, rendering the error of many-body quantum evolution upon truncation systematically computable. As clearly seen, the error stems entirely from interactions. In other words, the error does not depend on the choice of t_{ij} and even on the single particle energy matrix \epsilon_{ij} for any given truncated initial state, provided the evolution of the state, \partial_t |\Phi\rangle, is optimally taken as in Eq. , , . ### Determining the number of orbitals dynamically When the error becomes large or, alternatively, when we aim at describing the system more precisely, we have to increase the number of orbitals M_1 into M. The (M - M_1) additional orbitals then can be determined by variation of the error with respect to \phi_{u} (\vec{x},t) where M_1 < u \leq M, and subject to the orthonormalization condition \int \mathrm{d}\vec{x} ~ \phi_{u}^{*} (\vec{x},t) \phi_{v} (\vec{x},t) = \delta_{uv}. The method of Lagrange multipliers for functional variables gives the stationarity condition \begin{split} &\sum_{i,j,k,n,r,s,p,q=1}^{M_1} V_{ijku} \big( \hat{V}_{sp} \phi_{q} (\vec{x},t) \big) \langle \hat{a}_{ijkn}^{\dagger\dagger} \rangle \langle \rho \rangle_{nr}^{-1} \langle \hat{a}_{rspq}^{\dagger\dagger} \rangle \\ &- \sum_{i,j,p,q=1}^{M_1} \sum_{k,s=1}^{M} V_{ijku} \big( \hat{V}_{sp} \phi_{q} (\vec{x},t) \big) \langle \hat{a}_{ijk}^{\dagger\dagger} \hat{a}_{spq}^{\dagger} \rangle \\ &= \sum_{v=1}^{M} \mu_{uv} \phi_{v} (\vec{x},t) . \end{split} While this is an equation for the additional orbitals, the additional orbitals are difficult to obtain directly from the above equation. The best we can do is to guess some orbitals that will approximately satisfy the above equation. Therefore, we use the method of steepest constrained descent again. From an initial trial orbital, we propagate the orbital toward \begin{split} &\frac{d \phi_u (\vec{x})}{d \tau} = -\Delta_{\phi_u} (\tau) \bigg[ \sum_{i,j,k,n,r,s,p,q=1}^{M_1} V_{ijku} \big( \hat{V}_{sp} \phi_{q} (\vec{x},t) \big) \\ &\times \langle \hat{a}_{ijkn}^{\dagger\dagger} \rangle \langle \rho \rangle_{nr}^{-1} \langle \hat{a}_{rspq}^{\dagger\dagger} \rangle - \sum_{i,j,p,q=1}^{M_1} \sum_{k,s=1}^{M} V_{ijku} \big( \hat{V}_{sp} \phi_{q} (\vec{x},t) \big) \\ &\langle \hat{a}_{ijk}^{\dagger\dagger} \hat{a}_{spq}^{\dagger} \rangle - \sum_{v=1}^{M} \mu_{uv} \phi_{v} (\vec{x},t) \bigg] \end{split} so that the error become minimized with this additional orbital. If we propagate orbitals separately and iteratively one after another, i.e. only the uth orbital changes along \tau, the constraint becomes \int \mathrm{d}\vec{x}~ \phi_v^* (\vec{x}) \frac{\phi_u (\vec{x})}{d\tau} = 0 for v \neq u and \int \mathrm{d}\vec{x}~ \big( \phi_u^* (\vec{x}) \frac{\phi_u (\vec{x})}{d\tau} + \frac{\phi_u^* (\vec{x})}{d\tau} \phi_u (\vec{x}) \big) = 0. Then the undetermined Lagrange multipliers become \begin{split} \mu_{uv} = &\sum_{i,j,k,n,r,s,p,q=1}^{M_1} V_{ijku} V_{vspq} \langle \hat{a}_{ijkn}^{\dagger\dagger} \rangle \langle \rho \rangle_{nr}^{-1} \langle \hat{a}_{rspq}^{\dagger\dagger} \rangle \\& - \sum_{i,j,p,q=1}^{M_1} \sum_{k,s=1}^{M} V_{ijku} V_{vspq} \langle \hat{a}_{ijk}^{\dagger\dagger} \hat{a}_{spq}^{\dagger} \rangle , \end{split} where the orbital index ranges are constrained to be M_1 < u \leq M and 1 \leq v \leq M. %Explicit expression for the error in terms of correlation functions/orbitals/expansion coefficients? Since the inner rotation can be arbitrarily chosen, we can take the unit matrix, t_{ij}=\epsilon_{ij}, which renders the result in a simple form. The evolution Eq.  for the expansion coefficients becomes i \partial_t C_{\vec{n}} = \left\langle \vec{n} \right| \frac{1}{2} \sum_{i,j=1}^{M} \sum_{k,l=1}^{M_1} V_{ijkl} \hat{a}_{ijkl}^{\dagger\dagger} \left| \Phi \right\rangle . This implies the desired property that, when the interaction is turned off, the coefficients C_{\vec{n}} do not change at all. The Schr\ddot{o}dinger equation for the orbitals, Eq. , becomes \begin{split} &i \partial_t \phi_k (\vec{x},t) = \hat{h} \phi_k (\vec{x},t) \\ &+ \sum_{l=M+1}^{\infty} \sum_{i,n,p,q=1}^{M_1} V_{lnpq} \langle \rho \rangle_{ki}^{-1} \langle \hat{a}_{inpq}^{\dagger\dagger} \rangle \phi_l (\vec{x},t) \end{split} for k \leq M_1. The projection toward the outside of the sub-Hilbert space \mathcal{M} takes place only on the interaction term. Thus the option of taking t_{ij}=\epsilon_{ij} for the inner rotation shows the effect of the interaction term in this explicit manner. For M_1 < k \leq M, the evolution of the additional orbitals can be chosen in any convenient way, since the time derivatives of the additional orbitals do not occur in the error measure. The only constraint is the inner rotation. As we have chosen t_{lk} \big( \equiv \int \mathrm{d}\vec{x} \phi_l^* (\vec{x},t) i \partial_t \phi_k (\vec{x},t) \big) for M_1 < l \leq M and k \leq M_1 to be \epsilon_{lk}, t_{kl} should be t_{lk}^* = \epsilon_{lk}^* = \epsilon_{kl} to satisfy the orthonormality constraint Eq. . Taking the rotation toward the outside of the sub-Hilbert space to be also equal to the single-particle energy matrix, \epsilon_{kl} for k>M, the evolution of the additional orbital \phi_l (\vec{x},t) for M_1<l\leq M is determined by the simple equation i \partial_t \phi_l (\vec{x},t) = \hat{h} \phi_l (\vec{x},t), while the additional orbitals are found from Eq. . We, finally, emphasize again that the key difference from the MCTDHB approach is that additional macroscopically occupied orbitals during dynamical evolution can be found in our approach. By this means, we can handle the exceptional case when the SPDM is not invertible, and increase the number of orbitals under any given circumstances and boundary conditions. ## Numerical Implementation

Schematic potential we used. It is an infinite potential with width L and a finite bump with width a and height h(t) on center of the infinite potential. The ratio a/L is chosen to be 0.1 in the following implementations.

Here we present simple numerical implementations of our method. For the simplicity of the implementation, we investigate a system under infinite well potential. Then for the description of the orbitals, we can use DVR(discrete variable representation) with the finite number \bar{M} of sinusoidal basis functions. The i-th orbital \phi_i is described by the finite number of complex variables U_{\bar{k}i} through \phi_i (\vec{x},t) = \sum_{\bar{k}=1}^{\bar{M}} \phi_{\bar{k}} (\vec{x}) U_{\bar{k}i} (t) \quad\text{with}\quad \phi_{\bar{k}} (\vec{x}) = \sqrt{\frac{2}{L}} \sin \Big( \frac{\bar{k} \pi}{L} x \Big) where L is the width of the infinite well potential. Then the differentiations and integrations over the orbitals will be represented by matrices. The single-particle matrix elements are \begin{split} \epsilon_{ij} &= \int \mathrm{d} \vec{x} ~~ \phi_i^* (\vec{x}) \bigg[ -\frac{\nabla^2}{2m} + V_{\rm trap} (\vec{x}) \bigg] \phi_j (\vec{x}) \\ &= \sum_{\bar{k}, \bar{l}=1}^{\bar{M}} U_{i\bar{k}} \epsilon_{\bar{k}\bar{l}} U_{\bar{l}j} \end{split} where \epsilon_{\bar{k}\bar{l}} = \frac{\pi^2 \bar{k}^2}{2m L^2} \delta_{\bar{k}\bar{l}} + \int \mathrm{d} \vec{x} ~ \phi_{\bar{k}}^* (\vec{x}) V_{\rm trap} (\vec{x}) \phi_{\bar{l}} (\vec{x}) . When the trap potential is varied only by a overall factor along the time, i.e.  V_{\rm trap} (\vec{x},t) = h(t) V_{\rm trap} (\vec{x}), integrations which cost a bunch of time in calculation are needed just once in the beginning and the saved integration datas can be used afterwards saving lots of time. The two-body interaction elements are also calculated by \begin{split} V_{ijkl} &= \iint \mathrm{d} \vec{x}_{\alpha} \mathrm{d} \vec{x}_{\beta} ~ \phi_i^* (\vec{x}_\alpha) \phi_j^* (\vec{x}_\beta) V(\vec{x}_\alpha, \vec{x}_\beta) \phi_k (\vec{x}_\beta) \phi_l (\vec{x}_\alpha) \\ &= \sum_{\bar{p}, \bar{q}, \bar{r}, \bar{s}=1}^{\bar{M}} U_{i\bar{p}} U_{j\bar{q}} V_{\bar{p}\bar{q}\bar{r}\bar{s}} U_{\bar{r}k} U_{\bar{s}l} \end{split} where, when the interaction is given by V(\vec{x}_\alpha, \vec{x}_\beta) = g \delta(\vec{x}_\alpha - \vec{x}_\beta) , \begin{split} V_{\bar{p}\bar{q}\bar{r}\bar{s}} &= g \Big( \frac{2}{L} \Big)^2 \int_{0}^{L} \mathrm{d} x ~ \sin \Big( \frac{\bar{p} \pi}{L} x \Big) \sin \Big( \frac{\bar{q} \pi}{L} x \Big) \\ &~~~~~~~~~~~~~~~~~~~~~~~~ \times \sin \Big( \frac{\bar{r} \pi}{L} x \Big) \sin \Big( \frac{\bar{s} \pi}{L} x \Big) \\ &\!\!\!\!\!\!\!\!\!\!\!\!\!= \frac{g}{2L} \big( \delta_{\bar{p},-\bar{q}-\bar{r}-\bar{s}} -\delta_{\bar{p},\bar{q}-\bar{r}-\bar{s}} -\delta_{\bar{p},-\bar{q}+\bar{r}-\bar{s}} -\delta_{\bar{p},-\bar{q}-\bar{r}+\bar{s}} \\ &~~~ +\delta_{\bar{p},-\bar{q}+\bar{r}+\bar{s}} +\delta_{\bar{p},\bar{q}-\bar{r}+\bar{s}} +\delta_{\bar{p},\bar{q}+\bar{r}-\bar{s}} -\delta_{\bar{p},\bar{q}+\bar{r}+\bar{s}} \big) . \end{split}

Natural occupation number of the ground state on the first symmetric orbital according to the height h of the bump on the center of the infinite potential. Here we set L=10 \mu m, g=0.01, N=100. As h increases, particles get fragmented and gradually condensed from the symmetric orbital into the antisymmetric orbital. The details are discussed in the script.

Ground state without any bump. Therefore h=0, \hbar = m = 1 unit, and N=100, L=10 \mu m, g=0.01. Red dot indicates real value and blue squre indicates imaginary value of complex number. Green x in orbital graphs indicates |phi (x)|^2, i.e. square norm of the complex orbital at that position. The graph shows that the bosonic particles in the ground state are nearly condensated on the first symmetric orbital in (a). Since there is unitary transformation freedom or arbitrariness, the state can be represented in several different ways with different orbital bases. Consequently (b) is unitarilly transformed one from (a) and therefore these two states are exactly the same many-body state.

To simplify the calculations, we used \hbar = m = 1 unit and length in \mum unit. Therefore time is expressed in units of \frac{mL_{\mu m}^2}{\hbar}, and energy in units of \frac{\hbar^2}{m L_{\mu m}^2}. For example, \frac{mL_{\mu m}^2}{\hbar} = 1.37 msec and \frac{\hbar^2}{m L_{\mu m}^2} = 7.70 \times 10^{-32} \text{ J} = 2\pi \times 116 Hz when m is put to be the mass of ^{87}Rb atom . ### Ground state Using these numerical techniques and the method described in section , we firstly found the many-body ground state under different conditions. Since we are interested in the instantaneous error measure indicating the validity of the simulation, we tried to find two distinguished ground states which can be connected through the time evolution. Then we checked the ground state without any bump and with a bump on the center of the potential, expecting the condensation and the simpleset two-fold fragmentation of double well potential.

Ground state with a bump of height h=1000. Although (b) representation is more physically understandable, (a) represents the natural orbitals of eigenvalues of single particle density matrix. Here we used \hbar = m = 1 unit, and N=100, L=10 \mu m, g=0.01.

Ground state with a bump of height h=7000. Although (b) representation is more physically understandable, (a) represents the natural orbitals of eigenvalues of single particle density matrix. Here we used \hbar = m = 1 unit, and N=100, L=10 \mu m, g=0.01.

Fig.  shows the schematic potential we implemented. Raising the barrier height h, we figured out how the ground state and the degree of fragmentation changes. Fig.  shows how the fragmentation changes with respect to the height h of the bump. And fig. , , ,  shows several ground states obtained numerically with M=2 orbitals. Without any bumps, i.e. when h=0, the ground state is almost single condenstate as it is shown in (a) of fig. . 99\% of the particles are condensed on the first symmetric sinusoidal-like orbital. Just small excitation on the second orbital is there. But this state can be also represented as the coherent state with unitary transformation of the basis orbitals. See (b) in fig. . With left and right orbitals, we can see the state in a different view. This becomes quite useful in the following arguements analyzing the fragmentation. Seeing the change of fragmentation degree with respect to the height h in fig. , we noticed that the result is somewhat counter intuitive. Up to h=1000, the condensation drops gradually. But between h=1000 and h=6000, there is a plateau. And from h=6000 to h=8000, the particles move from the first symmetric orbital to the second antisymmetric orbital. See (a)'s of fig. , , , . Though we expected that the condensation would be two-fold fragmented as h goes up, the result was different from it. %As the bump on the center of the potential goes up, i.e. h increases, the ground state becomes two-fold fragmented. To resolve or explain this irregularity, we checked what the states are looked like and how the state is changing. With seeing only the natural orbitals, which are presented in (a)'s of fig. , , , , and with which the single particle density matrix \langle \hat{a}^{\dagger}_{ij} \rangle is diagonalized, it was hard to understand the states physically. When C_{\vec{n}}'s are widely distributed and therefore various features are mixed and superposed, the conceptual understanding of the state becomes quite obscure. Then we transformed the orbital bases unitarilly to see the same state in a different view and understand it better. See (b)'s in fig. , , , . As it is shown, the distribution of C_{\vec{n}} is quite sharp around center even when the degree of fragmentation is around 85\% and 15\%, i.e. when judged from the eigenvalues of SPDM, the state is regarded as being quite condensated. From these features, we can conclude or argue that this peculiarity is due to the lack of orbitals, i.e. because too small number of orbitals, in this case M=2, are used.

Ground state with a bump of height h=11000. Here we used 1hbar = m = 1 unit, and N=100, L=10 \mu m, g=0.01.

For example, we need at least M=4 orbitals to describe two independent, in other words spatially quite separated, condensates when we need at least M=2 orbitals to describe each condensate well enough. In extreme condition of our case where h \rightarrow \infty, left and right are detangled. And therefore, in this case, we might need M=4 orbitals to represent the ground state properly. Since the ground state on each left and right infinite well potential is nearly single condensate which is similar to the one in fig. , M=2 can be enough in this extreme case. But as our cases implemented are in between, surely it can be insisted that at least M=4 orbitals are needed to describe the ground state quite properly. In conclusion, the non fragmented weird feature above h \approx 2000 can be regarded as a side effect of truncating too many orbitals or information about the state, which is generally done to save costs and time of simulation. Since the small excitation on each left and right potential, i.e. excitation on the third and fourth orbital, cannot be counted with M=2 simulation, these features might be compensated by the mixing of C_{\vec{n}} around n_1 = n_2 = 50. Consequently this causes the decrease of the degree of fragmentation. Although the more concrete argument requires M=4 simulations, it also requires exponentially increased costs, efforts and time. For now, seeing the unitarily transformed states and handling the datas delicately, we claim that, even with M=2 simulation, many interesting features can be seen enough and be trustworthy. And for the goal of checking the validity of simulation of time evolution through the instantaneous error measure, this example with M=2 is just appropriate. %\colb{\% This is comment: For more complete paper or confirmation on this argument, we gonna need an additional simulations with at least M=4 orbitals. But this will take too long and require lots of efforts additionally.} ### Time evolution Expecting dynamical fragmentation of single condensate into two-fold fragmented state, we ramped up the middle of the infinite well potential . With M=1 simulation, the dynamical fragmentation cannot be described. So we can expect that the indicator or the instantaneous error of M=1 evolution would show some increases at the point where the fragmentation emerges whereas the error of M=2 does not show any increment.

(a) the instantaneous error Eq.  of M=1 simulation. (b,c) particle occupation along the time in different y-scale.

Seeing Fig. , we can see the correspondence between the error and the fragmentation. But there is also troublesome offset problem in this error measure. To resolve this problem, we see meaning of the error in a different view. ### Physical meaning of the error The one crucial hypothesis in our method is that the most part of the bosons of our interest resides only within M-orbitals during the time of our interest. In mathematical words, it says \langle \sum_{i=1}^{M} \hat{n}_{i} (t) \rangle \approx N \quad \text{when } t<T_{\text{interest}}. Expanding it with Taylor series, we can see some short time tendancy of the number of particles in M-orbitals \langle \sum_{i=1}^{M} \hat{n}_{i} (t) \rangle = N + \frac{d \langle \sum_{i=1}^{M} \hat{n}_{i} (t) \rangle}{d t} t + \frac{d^2 \langle \sum_{i=1}^{M} \hat{n}_{i} (t) \rangle}{2 d t^2} t^2 + \cdots . Since the initial many-body state consists of only M-orbitals or less than M-orbitals (at least we can assume that the whole particles are initially within M-orbitals) and the commutation relation satisfies [\sum_{i=1}^{\infty} \hat{n}_i , \hat{H}] = 0, the first derivative \begin{split} &i \frac{d}{d t} \langle \sum_{i=1}^{M} \hat{n}_{i} (t) \rangle \\ &= \langle \sum_{i=1}^{M} \hat{n}_{i} (t) \hat{H} - \hat{H} \sum_{i=1}^{M} \hat{n}_{i} (t) + \sum_{k=1}^{\infty} \sum_{i=1}^{M} ( t_{ki} \hat{a}^{\dagger}_{ki} - t_{ik} \hat{a}^{\dagger}_{ik} ) \rangle \end{split} becomes automatically zero. Then it means that the total occupation number on M-orbitals doesn't change instantly if the initial state has all bosons under M-orbitals. Seeing the second derivatives of number conserving in M-orbitals, \begin{split} &\frac{d^2}{d t^2} \langle \sum_{i=1}^{M} \hat{n}_{i} (t) \rangle = -2 \langle \hat{H} \Big( N - \sum_{i=1}^{M} \hat{n}_{i} \Big) \hat{H} \rangle - \langle \big[ \partial_t^2 \sum_{i=1}^{M} \hat{n}_{i} (t) \big] \rangle \\ &+ 2 \langle \hat{H} \big[ i \partial_t \sum_{i=1}^{M} \hat{n}_{i} (t) \big] \rangle - 2 \langle \big[ i \partial_t \sum_{i=1}^{M} \hat{n}_{i} (t) \big] \hat{H} \rangle . \end{split} Then putting the change of orbitals given in Eq.  and arranging the equation, it gives \begin{split} -\frac{d^2}{d t^2} \langle \sum_{i=1}^{M} \hat{n}_{i} (t) \rangle = &2 \sum_{l=M+1}^{\infty} \Big[ \sum_{i,j,p,q=1}^{M_1} \sum_{k=1}^{\infty} V_{ijkl} V_{lkpq} \langle \hat{a}_{ijpq}^{\dagger\dagger} \rangle \\ &+ \sum_{i,j,k,s,p,q=1}^{M_1} V_{ijkl} V_{lspq} \langle \hat{a}_{ijskpq}^{\dagger\dagger\dagger} \rangle - \sum_{i,j,k,n,r,s,p,q=1}^{M_1} V_{ijkl} V_{lspq} \langle \hat{a}_{ijkn}^{\dagger\dagger} \rangle \langle \rho \rangle^{-1}_{nr} \langle \hat{a}_{rspq}^{\dagger\dagger} \rangle \Big] . \end{split} This has quite similar form to the instantaneous error of Eq.  since it has double derivative with repect to time and therefore \hat{H}^2 term. The leaking bosons from M-orbitals then can be approximately counted to be \langle \sum_{i=1}^{M} \hat{n}_{i} (t) \rangle \approx N + \frac{d^2 \langle \sum_{i=1}^{M} \hat{n}_{i} (t) \rangle}{2 d t^2} t^2 . Therefore our method with M-orbitals will be valid only for the time scale T which satisfies - \frac{d^2 \langle \sum_{i=1}^{M} \hat{n}_{i} (t) \rangle}{2 d t^2} T^2 \ll N . This means T \ll \sqrt{ \frac{2N}{-d^2 \langle \sum_{i=1}^{M} \hat{n}_{i} (t) \rangle / d t^2} } where the many-body state in this equation is put by the initial one and the instantaneous state afterwards. The upper bound of time will be updated continuously, using the instantaneous state. We have checked this double derivative term. Comparing the values of the error and double derivative of the particle conservation, we can see the value of the d_t^2 is just 4 times of the error. \col{But this factor 4 is only when the term below is dominant? You found this numerically to be true or analytically?} From this hint, we have found that the term V_{ijkl} V_{lkpq} \langle \hat{a}_{ijpq}^{\dagger\dagger} \rangle is the most dominant one both in Eq.  and Eq. . Even though Eq.  does not exactly correspond to Eq. , we can adapt some physical concepts of Eq.  into the instantaneous error and estimate valid time scale from it.

(Graph will be put.) The error divided by double derivative of the particle conservation.

\label{error vs dt2}

But still the big offset is troublesome. In Fig. , the offset of the error is almost 10 times bigger than the change of the error along time. This might be because the small fraction outside of M-orbitals is not appropriately counted. Actually we ignored that part as we set n_k = 0 for k>M orbitals along all time. The revived small part will compensate the flow from inside M to outside of M-orbitals by the flow from outside to inside of M. Since this flow will be nearly constant if the small fraction stays small steadily along time, we can argue that we can hopefully ignore the offset and concentrate on the change of the error along time. This statement is also supported by the fact that the offset value of M=2 simulation is significantly diminished when compared to M=1 simulation. \col{Obviously, this crucial argument has to be refined still, as it is not yet accurate. What do you mean by ``revived'' small part?} Therefore seeing Fig.  and concentrating on the change of the error (\approx 20 \times 10^{-3}), we can see the valid time scale Eq.  (\approx 100) is consistant to the time scale where the fragmentation occurs. Tracking the change of the error, we can actually sense the validity of the simulation. ## Summary \colb{Theory is estabilished by observing and analyzing experiments about physical systems. And the theory is continuously verified and modified through successive experiments and endless analyses. Sometimes the theory expect first some interesting and useful results, and the experiments follow the expectation. But the equation of the theory is not always exactly solvable, so we approximate the equation within a plausible form by fair reasonings to solve the problem. In these processes, when the well-designed experiment does not agree with the theoretical anticipation, we have to check whether the theory is wrong and therefore needs to be modified, or the approximate treatment of the theory was too harsh and more accurate calculation is needed. But it will be much better if we can assert that the approximated simulation is quite the same to the exact manipulation of the theory before a direct comparison with experimental outputs. Since there can be mistakes in experiment too, discrimination of theoretical mistakes from experimental one is crucial to establish a sound theory.} \col{What is the point of this very philosophical paragraph? I do not quite see the relation to our paper. It is more important to rewrite what comes below, and incorporate the discussion on the numerical implementation and its outcome there. We have no experiment in our paper, so please rewrite the above such that there is a connection to our paper made.} Using McLachlan's principle and the methods of Lagrange multipliers and steepest constrained descent, we have found a computationally feasible way to describe the many-body evolution of bosons in a rigorously controlled manner. Writing the many-body state in Hartree form and limiting the size of the Hilbert space by truncating into a finite number of macroscopically occupied field operator modes, the error from the exact evolution can be minimized self-consistently. This gives a variationally optimized solution to the evolution of the truncated many-body state. We have demonstrated that without two-body interactions, our scheme possesses the desired property that the evolution of the many-body state can be exactly described with zero error, cf.  Eq. . When interactions are turned on, the error accumulates during time evolution. Employing our method, we can evolve the truncated many-body state in an optimized way. Monitoring the error simultaneously, we can increase the accuracy of the evolution by increasing the number of orbitals in a self-consistent way. By adaptively changing the number of orbitals based on the instantaneous error measure, we can automatically ensure the validity of the result for the many-body state. We conclude by a brief summary of our approach applied to the well-known and ubiquitous Gross-Pitaevski\check{i} equation. We start by evolving the initial trial state |\Phi\rangle along the M=1 version of Eq. ,  and simultaneously check whether the error Eq.  remains small or not. Monitoring the error Eq. , we can determine under which conditions the Gross-Pitaevski\check{i} equation loses its validity. When this happens, the error becoming large, we increase the number of orbitals to M=2 (thus, here, M_1=1 and M=2). The additional second orbital is found by the method of steepest constrained descent, using Eq. . Then the subsequent evolution of the quantum many-body state follows the M=2 version of Eq. , , while the evolution of the (initially singular) second orbital follows Eq. . We monitor the error Eq.  again, checking that the error is decreased to a sufficient degree. In a self-consistent manner the processes described are carried out successively until we obtain a prescribed accuracy. ## RRA

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18. We illustrate this statement as follows. The reason why we can use df = \frac{\partial f}{\partial z}dz + \frac{\partial f}{\partial z^*}dz^* is not because dz and dz^* are linearly independent differentials, but because dz and dz^* are differentials, that is infinitesimal quantities; dz^* is in fact 100\% dependent on dz since dz^* is simply the complex conjugate of dz. Therefore we cannot simply argue that because dz and dz^* are independent variations, the partial derivatives \frac{\partial f}{\partial z} and \frac{\partial f}{\partial z^*} must be zero for the stationarity of f. There is a rather widespread misunderstanding of this mathematical fact found in the literature.
19. When the variables and constraints are given in complex form, the method of Lagrange multipliers is expressed by \frac{\partial f} {\partial z_k^*} = \sum_l \bigg[ \lambda_l \frac{\partial g_l} {\partial z_k^*} + \lambda_l^* \frac{\partial g_l^*} {\partial z_k^*} \bigg]. Though g_l (z_k, z_k^*)=c and g_l^* (z_k, z_k^*)=c^* are the same constraint, we have to add the complex conjugate of it in the Lagrange equation. Be careful to note that (\frac{\partial g_l}{\partial z_k^*})^* is not equal to \frac{\partial g_l^*}{\partial z_k^*}. The above equation gives the condition for the stationarity of the real-valued function f.
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25. kipid's blog - Method of Lagrange multipliers
26. Second quantization of many-body Quantum mechanics