- Open Access
A dynamic systems approach to fermions and their relation to spins
© Zimborás et al.; licensee Springer on behalf of EPJ. 2014
Received: 8 April 2014
Accepted: 21 July 2014
Published: 11 September 2014
The key dynamic properties of fermionic systems, like controllability, reachability, and simulability, are investigated in a general Lie-theoretical frame for quantum systems theory. It just requires knowing drift and control Hamiltonians of an experimental set-up. Then one can easily determine all the states that can be reached from any given initial state. Likewise all the quantum operations that can be simulated with a given set-up can be identified. Observing the parity superselection rule, we treat the fully controllable and quasifree cases of fermions, as well as various translation-invariant and particle-number conserving cases. We determine the respective dynamic system Lie algebras to express reachable sets of pure (and mixed) states by explicit orbit manifolds.
PACS Codes: 03.67.Ac, 02.30.Yy, 75.10.Pq.
The vast experimental progress in implementing coherent control of ultra-cold gases including fermionic systems [1–6] has also great impact on quantum simulation (e.g., ) of quantum phase transitions [8, 9], pairing phenomena , and in particular for understanding phases in Hubbard models . Moreover, digital quantum simulation of fermionic systems has come into focus [12–16]. For either way of quantum simulation, there are important algebraic aspects going beyond the standard textbook approach , some of which can be found in [18–21]. Here we set out for a unified picture of quantum systems theory in a Lie-algebraic frame following the lines of . It paves the way for optimal-control methods to be applied to fermionic systems and leads to a plethora of new results presented here.
It is generally recognized that optimal control algorithms are key tools needed for further advances in experimentally exploiting these quantum systems for simulation as well as for computation [23–26]. In the implementation of these algorithms it is crucial to know before-hand to which extent the system can be controlled. For instance: which states can be reached from a given initial state under given controls? or likewise: which quantum operations can be simulated in a given set-up? The usual scenario (in coherent control) is that one is given a drift Hamiltonian and a set of control Hamiltonians with tunable strengths. The achievable operations will be characterized by their generators forming the system Lie algebra. Then the reachable sets of states can easily be given as the respective state orbits under the corresponding dynamic group. Dynamic Lie algebras and reachability questions have been intensively studied in the literature for qudit systems [22, 27–29]. However, in the case of fermions these questions have to be reconsidered mainly due to the presence of the parity superselection rule. Hence in a broader sense the present work on fermions can be envisaged also as a step towards quantum control theory for quantum simulation in the presence of superselection rules.
Apart from discussing the implications of the parity superselection rule we treat the cases of imposing translation-invariance or particle-number conservation. In particular, the experimentally relevant case of quasifree fermions (with and without translation invariance) is discussed in detail. Since we interrelate fermionic systems with the Lie-theoretical framework of quantum-dynamical systems, at times we will be somewhat more explicit and put known results into a new frame. The main results thus extend from general fermionic systems to the action of Hamiltonians with and without restrictions like quadratic interactions, translation invariance, reflection symmetry, or particle-number conservation.
The paper itself is structured as follows: In order to set a unified frame, we resume some basic concepts of Hamiltonian controllability of qudit systems in Section 2. Thus the dynamic systems approach is presented in a way to address a broader readership, who is enabled to make quick use of the key results summarized in the tables. These concepts are subsequently translated to their fermionic counterparts, starting with the discussion of general fermionic systems in Section 3.
Then the new results are presented in the following six sections: In Section 4 we obtain the dynamic system algebra for general fermionic systems respecting the parity superselection rule (see Theorem 4 in Section 4.1). An explicit example for a set of Hamiltonians that provides full controllability over the fermionic system is discussed in Section 4.2. Some general results on the controllability of fermionic and spin systems, such as Theorem 51, are relegated to Appendix A. Following the same line, in Section 5 we wrap up some known results on quasifree fermionic systems in a general Lie-theoretic frame by streamlining the derivation for the respective system algebra in Proposition 9 of Section 5. Corollary 16 provides a most general controllability condition of quasifree fermionic systems building on the tensor-square representation used in . Furthermore, we develop methods for restricting the set of possible system algebras by analyzing their rank, see Theorem 13 as well as Appendices C and D. The structure and orbits of pure states in quasifree fermionic systems are analyzed in Section 6 leading to a complete characterization of pure-state controllability (Theorem 23). Sections 7 and 8 are devoted to translation-invariant systems. For spin chains we give in Theorem 25 the first full characterization of the corresponding system algebras and strengthen in Theorem 27 earlier results on short-range interactions in . The system algebras for general translation-invariant fermionic chains are given in Theorem 30 of Section 7.3. We also identify translation-invariant fermionic Hamiltonians of bounded interaction length which cannot be generated from nearest-neighbor ones (see Theorem 33 of Section 7.4). The particular case of quadratic interactions (see Section 8.1) is settled in Theorem 34. Corollary 35 considers systems which additionally carry a twisted reflection symmetry (or equivalently have no imaginary hopping terms) as discussed in . Furthermore, we provide a complete classification of all pure quasifree state orbits in Theorem 39 of Section 8.2. This leads to Theorem 41 of Section 8.3 presenting a bound on the scaling of the gap for a class of quadratic Hamiltonians which are translation-invariant. Section 9 deals with fermionic systems conserving the number of particles. Their system algebras in the general case as well as in the quasifree case are derived in Proposition 42 and Proposition 43, respectively. Furthermore, a necessary and sufficient condition for quasifree pure-state controllability in the particle-number conserving setting is provided by Theorem 48.
In Section 10, we summarize the main results as given in Theorem 4, Corollary 16, as well as in Theorems 23, 25, 27, 30, 33, 34, 39, 41, and 48. We conclude leaving a number of details and proofs to the Appendices in order to streamline the presentation.
2 Basic quantum systems theory of N-level systems
driven by the Hamiltonian and fulfilling the initial condition . Here the drift term describes the evolution of the unperturbed system, while the control terms represent coherent manipulations from outside. Equation (1) defines a bilinear control system Σ , as it is linear both in the density operator and in the control amplitudes .
saying that a N-level quantum system is fully controllable if and only if its system algebra is the full unitary algebra, which we will relax to in a moment. This notion of controllability is also intuitive (recalling that the exponential map is surjective for compact connected Lie groups), as it requires that all Hamiltonian directions can be generated.
So in fully controllable closed systems, to every initial state the reachable set is the entire unitary orbit . With density operators being Hermitian, this means any final state can be reached from any initial state as long as both of them share the same spectrum of eigenvalues (including multiplicities). Thus the reachable set of is the isospectral set of .
and is a proper subgroup of .
In , we have analyzed the possibility of quantum simulation with respect to the dynamic degrees of freedom and have given a number of illustrating worked examples.
More generally, let denote the commutant of a set S of matrices, i.e., the set of all complex matrices which commute simultaneously with all matrices in S. By Jacobi’s identity one gets two properties of the centralizer pertinent for our context: First, an element s that commutes with all Hamiltonians also commutes with their Lie closure (i.e. ), as and imply . Second, for any , and imply , so the centralizer forms itself a Lie subalgebra to consisting of all symmetry operators.
where denotes the real span. Clearly, generates the stabilizer group to the state space of the control system .
following from Eq. (1), one may take equally well from or . Thus henceforth we will only consider special unitaries (of determinant +1) generated by traceless Hamiltonians , since for any Hamiltonian there exists an equivalent unique traceless Hamiltonian generating a time evolution coinciding with the one of .
However, the above simple arguments are in fact much stronger, e.g., one readily gets the following statement:
is equivalent to .
Proof Using the formula we show that Eq. (11) is equivalent to condition (a): for all non-negative integer k and all . Moreover, (a) implies condition (b): for all non-negative integer k and all , as . Also, (a) follows from (b) due to . Applying to (b) completes the proof. □
Therefore, let us consider a pair of Hamiltonians (fulfilling the conditions of Lemma 2) as equivalent on the state space , if their difference falls into the centralizer .
Finally note that all unitary conjugations of type are elements of the projective special unitary group , where the centers of and are respectively given by and . Moreover, recall , where can be represented as commutator superoperator . Now, for any , one immediately obtains , which also elucidates that the generators of the projective unitaries are given by .
3 Fermionic quantum systems
In this section, we fix our notation by recalling basic notions for fermionic systems. In the first subsection, we discuss the Fock space and different operators acting on it as given by the creation and annihilation operators as well as the Majorana operators. We point out how the Lie algebra of skew-Hermitian matrices can be embedded as a real subspace in the set of the complex operators acting on the Fock space. In the second subsection, we focus on the parity superselection rule and how it structures a fermionic system.
3.1 The Fock space and Majorana monomials
Given an orthonormal basis of , the Fock vacuum and the vectors of the form (with and ) form an orthonormal basis of . Note that is a -dimensional Hilbert space isomorphic to ().
where denotes the anticommutator. Moreover, every linear operator acting on can be written as a complex polynomial in the creation and annihilation operators.
where the following notation for the Pauli matrices , , and is used.
Similarly, one obtains a basis of by excluding .
3.2 Parity superselection rule
An additional fundamental ingredient in describing fermionic systems is the parity superselection rule. Superselection rules were originally introduced by Wick, Wightman, and Wigner  (see also [40, 41]). These rules, in the finite-dimensional definition of Piron , describe the existence of non-trivial observables that commute with all physical observables. The existence of such a commuting observable in turn implies that a superposition of pure states from different blocks of a block-diagonal decomposition w.r.t. the eigenspaces of this observable are equivalent to an incoherent classical mixture.
These physical operators are also exactly the ones that can be written as a sum of products of an even number of Majorana operators (as P contains all Majorana operators whereof there exist an even number). They are therefore denoted as even operators for short. If the parity is the only non-trivial symmetry, we obtain , where the bracket stands for the complex-linear span.
Note that for clarity we use this notation in contrast to the notation of even and odd subspaces (which is also used in the literature) in order to avoid any confusion with the even operators.
Hence physical observables cannot distinguish between the density operator ρ and its block-diagonal projection to (which is always an even density operator). In this sense, a physical linear combination (a formal superposition) of pure states from the positive and negative parity subspaces is equivalent to an incoherent classical mixture. Equation (15) also shows that without loss of generality we can restrict ourselves to even density operators and regard only those as physical.
Finally, we would like to recall three further aspects of the parity superselection rule. First, without the parity superselection rule, two noncommuting observables acting on two different and spatially-separated regions would exist which would allow for a violation of locality (e.g., by instantaneous signaling between the regions). Second, the parity superselection rule, of course, does not apply if one uses a spin system to simulate a fermionic system via the Jordan-Wigner transformation. This system respects locality, since the Majorana operators are—in this case—localized on the first spins; two non-commuting Majorana operators are therefore not acting on spatially-separated regions. Third, the parity superselection rule also affects the concept of entanglement as has been pointed out and studied in detail in [43, 44].
4 Fully controllable fermionic systems
the on-site potential of the first site , and a pairing-hopping term between the first two modes (see Proposition 6). Finally, we provide a general discussion about when the commutant of a system algebra determines the algebra itself.
4.1 System algebra
In the case of qubit systems mentioned in Section 2, two Hamiltonians generate equivalent time evolutions if and only if they differ by a multiple of the identity. This condition can readily be modified for the fermionic case such as to match the parity-superselection rule as well.
Corollary 3 Let and be two physical fermionic Hamiltonians, i.e., even Hermitian operators acting on . Then by Lemma 2 the equality holds for all even (physical) density operators with in the sense that and generate the same time-evolution, if and only if with .
and moreover, and H are equivalent and generate the same time evolution. If necessary, we can restrict ourselves to the set of Hamiltonians satisfying Eq. (16). These elements decompose as , where and are generic traceless Hermitian operators each acting on a -dimensional Hilbert space. We explicitly define the linear space of physical fermionic Hamiltonians as generated by the basis of all even Majorana monomials without the operators and P, ensuring that is traceless both on and .—We summarize our exposition on fully controllable fermionic systems in the following result:
Theorem 4 The Lie algebra corresponding to the physical fermionic (and Hermitian) Hamiltonians is . The most general set of unitary transformations generated by is given as the block-diagonal decomposition . Hence a set of Hermitian Hamiltonians defines a fully controllable fermionic system iff .
Remark 5 For Lie algebras, will denote only an abstract direct sum without referring to any concrete realization. We reserve the notation to specify a direct sum of Lie algebras which is (up to a change of basis) represented in a block-diagonal form .
Proof of Theorem 4 It follows from Section 3 that commutes with P and that the matrix representation of splits into two blocks of dimension corresponding to the + and − eigenspaces of P. As the center of is given by , the Lie algebra is semisimple. As there are exactly linear-independent operators in , the system algebra could be . And indeed, all other system algebras are ruled out as the subalgebras acting on each of the two matrix blocks would have a smaller Lie-algebra dimension than . □
4.2 Examples and discussion
We start out with an example realizing a fully controllable fermionic system by adding only one quartic operator to the set of quadratic Hamiltonians which will be discussed in Section 5 below (cf. Theorem 11):
Proof It follows from the independent Theorem 11 (see Section 5 below) that , , and generate all quadratic Majorana monomials . Consider an even Majorana monomial of degree , where is defined using the ordered index set ℐ, and a quadratic operator with and . We can change any index p of into q of using . Therefore, we get from and the quadratic operators all Majorana monomials of degree four.
Using the quartic Majorana monomials we can increase the degree of the monomials in steps of two: Consider the operators and which are defined using the ordered index sets ℐ and and have degrees and 4, respectively. Assuming that , we can generate an operator of degree where the corresponding ordered index set is given by . By induction, we can now generate all even Majorana monomials except . Note that cannot be obtained as holds by construction. Thus, we get all elements of (see Section 4.1) and the proposition follows. □
The proof also implies that all the operators generated commute with (cf. Eq. (14)) and the identity operator . In addition, all operators commuting simultaneously with all elements of can be written as a complex-linear combination of and P. We thus obtain a partial characterization of full controllability in fermionic systems:
Lemma 7 Consider a fermionic quantum system with modes. A necessary condition for full controllability of a given set of Hermitian Hamiltonians is that .
One can expect that the condition of Lemma 7 is not sufficient under any reasonable assumption by applying counterexamples from spin systems in . These counterexamples could be lifted to fermionic systems by providing the explicit form of the embeddings from to the first and second component of the direct sum .
We consider the algebras and , where the latter algebra is isomorphic to and contains the additional (non-physical) generator . Note that , i.e., the centralizers of both algebras are equal. However . In particular, we have , and does not fulfill the double-centralizer property. A more general incarnation of this effect in line with a discussion of double centralizers is given in Appendix A. It leads in the case of irreducible subalgebras to the following maximality result:
To sum up, the symmetry properties of a Lie algebra , as given by its commutant w.r.t. a representation of , do not determine the Lie algebra uniquely. Yet the commutant allows us to infer a unique maximal Lie algebra contained in , which is (up to an identity matrix) equal to the double commutant of , but in general not to itself.
5 Quasifree fermions
Here we present the dynamic system algebras for fermions with quadratic Hamiltonians. For illustration, also the relation to spin chains is worked out in detail. In this context, we show by free fermionic techniques that a Heisenberg-XX Hamiltonian of Eq. (21) combined with the one-site term and the two-site interaction gives rise to the system algebra (see Theorem 11), while the first two operators generate only the subalgebra (see Theorem 13). Further results along this line are presented in Appendix C.
Finally, we arrive at a very useful general result: In order to decide if a set of operators generates the full quadratic algebra for d modes, we characterize quadratic operators by a real skew-symmetric matrix T whose entries are given via (see Eq. (19)). Adapting our tensor-square criterion for full controllability from spin systems  to quasifree fermionic systems, a set of operators generates the full quadratic algebra if and only if the joint commutant of the operators has dimension three (see Corollary 16).
5.1 Quadratic Hamiltonians
where the coupling coefficients and are complex entries of the -matrices A and B, respectively. The canonical anticommutation relations and the hermiticity of H require that A is Hermitian and B is (complex) skew-symmetric. The terms corresponding to the non-zero matrix entries of A and B are usually referred to as hopping and pairing terms, respectively. Related parameterizations for quadratic Hamiltonians are discussed in Appendix B.
one can easily verify that the space of quadratic Hamiltonians is closed under the commutator. To sum up, we have established the well-known Lie homomorphism from the system algebra generated by a set of quadratic Hamiltonians (whose control functions are given by the matrix entries of A and B) onto the system algebra represented by the entries of T (cf. pp.183-184 of ):
Proposition 9 The maximal system algebra for a system of quasifree fermions with d modes is given by .
Proof Let the map h transform the Majorana monomial into the skew-symmetric matrix where has the matrix entries . We show that h is a Lie-homomorphism assuming and in the following, while the case of or holds trivially. Note that . It follows from Eq. (20b) that = = . □
5.2 Examples and explicit realizations
Lemma 10 Consider a fermionic quantum system with modes. The system algebras and generated by the set of Lie generators and contain the elements with for all as well as with and with for all .
Note that the elements , , and are mapped by the Jordan-Wigner transformation to the spin operators , , and , respectively.
Proof of Lemma 10 We compute the commutators , , and . We can now reduce the problem from d to by subtracting from . The cases of can be verified directly and the proof is completed by induction. □
This proof also yields an explicit realization for the algebra while providing a more direct line of reasoning as compared to our proof of Theorem 32 in .
Theorem 11 Consider a fermionic quantum system with modes. The system Lie algebra generated by is given by .
Proof The cases of can be verified directly. We build on Lemma 10 and remark that as it is generated only by quadratic operators (see Proposition 9). We compute in the Jordan-Wigner picture , and . This shows by induction that . As is a maximal subalgebra of (see p.219 of  or Section 8.4 of ), one obtains that . Alternatively, one can explicitly show that consists of all quadratic Majorana operators, which together with Proposition 9 also completes the proof. □
Note that the generators , , and can be described using the Hamiltonian of Eq. (18) while keeping the control functions given by the matrix entries and in the real range, see Appendix B for details. This also provides a simplified approach to Theorem 32 in , where only the real case was considered:
Corollary 12 (see Theorem 32 in )
Consider a control system given by the Hamiltonian components of Eq. (18). The control functions are specified by the matrix entries and which are assumed to be real. The resulting system algebra is .
The relations between quasifree fermions and spin systems will be analyzed in Appendix C.—Next we treat the case of the algebra .
Theorem 13 (see Lemma 36 in )
Consider a fermionic quantum system with modes. The system Lie algebra generated by is given by .
Here we just sketch ideas for the proof of Theorem 13 while leaving the full details to Appendix D. Our methods exploit the detailed structure of the appearing Majorana operators while being more explicit than in  and avoiding obstacles of the spin picture. Building on the notation of Lemma 10, we show that the elements with together with the elements with and with where and form a basis of . One obtains that . Furthermore, the elements form a maximal abelian subalgebra and the rank of is equal to d. (The rank of a Lie algebra is defined as the dimension of its maximal abelian subalgebras.) We limit the possible cases further by showing that is a direct sum of a simple and a one-dimensional Lie algebra. A complete enumeration of all possible cases completes the proof.
Remark 14 A spin chain equivalent to the fermionic system in Theorem 13 was also considered in , where it was shown how to swap pairs of fermions using the given Hamiltonians. As a consequence of Theorem 13, the Lie algebra in the spin chain of  can be identified as . Clearly, its size grows only linearly with the number of modes d. However, the addition of controlled-Z gates, as discussed in , already allows for scalable quantum computation.
5.3 Tensor-square criterion
Consider a control system of quasifree fermions which is represented by matrices in the form of Eq. (19). For more than two modes (i.e. ), we can efficiently decide if the system algebra is equal to . Recall that the alternating square and the symmetric square of a representation ϕ are defined as restrictions to the alternating and symmetric subspace of the tensor square .
Theorem 15 Assume that is a subalgebra of with and denote by Φ the standard representation of . Then, the following statements are equivalent: (1) . (2) The restriction of Alt2Φ to the subalgebra is irreducible and the restriction of Sym2Φ to splits into two irreducible components. Each irreducible component occurs only once. (3) The commutant of all complex matrices commuting with the tensor square of has dimension three.
Proof Assuming (1), condition (2) follows from the formulas for the alternating and symmetric square of with given in its standard representation [where denotes the corresponding highest weight]: The alternating square is given as for and for in the case of (cf. Table 6 in  or Table X in ). The symmetric square for and can be computed using Example 19.21 of . We verify the dimension of the commutant and show that (3) is a consequence of (2) by applying Proposition 50 which says that the dimension of the commutant of a representation ϕ is given by where the are the multiplicities of the irreducible components of ϕ. For the rest of the proof we assume that condition (3) holds. We remark that the representation is irreducible as otherwise the dimension of the commutant would be larger than three. Thus, we obtain that is semisimple. The dimension of the commutant allows only two possibilities: one of the restrictions or to the subalgebra has to be irreducible. We emphasize that is given in an orthogonal representation (i.e. a representation of real type) of even dimension, as is given in an irreducible representation obtained by restricting the standard representation of . Therefore, we can use the list of all irreducible representations which are orthogonal or symplectic (i.e. of quaternionic type) and whose alternating or symmetric square is irreducible (Theorem 4.5 as well as Tables 7a and 7b of ): (a) for the alternating square of the symplectic representation of dimension two, (b) for the alternating square of the orthogonal representation of dimension three, (c) for with the alternating square of the orthogonal representation of dimension , (d) for with the alternating square of the orthogonal representation of dimension 2ℓ, and (e) for with the symmetric square of the symplectic representation of dimension 2ℓ. Only possibility (d) fulfills all conditions which proves (1). □
Describing the matrices in the tensor square more explicitly along the lines of , we present a necessary and sufficient condition for full controllability in systems of quasifree fermions.
Corollary 16 Consider a set of matrices as given by Eq. (19) generating the system algebra with . We obtain if and only if the joint commutant of has dimension three.
Along the lines of Eq. (19), one can apply Corollary 16 to the matrices T corresponding to the generators of of Theorem 11. For one can verify that the commutant of the tensor square has dimension three. But for one computes a dimension of four as is not simple.
and thus holds jointly for all ; denotes the zero matrix of degree N.
6 Pure-state controllability for quasifree systems
In this section, we present a straightforward criterion for pure-state controllability of quasifree fermionic systems with d modes. A fermionic state is called quasifree if Majorana operators of odd degree map it to zero and even-degree ones map it to states which factorize into the Wick expansion form (see below). We obtain that quadratic Hamiltonians act transitively on pure quasifree states, i.e., every pure quasifree state can be transformed into any other pure quasifree state using only quadratic Hamiltonians (see Theorem 20).
In particular, within the Lie algebra of quadratic Hamiltonians a subalgebra isomorphic to provides the stabilizer of any pure quasifree state. Thus the set of pure quasifree states can be identified with a homogeneous space of the type . At first glance, this might suggest that for full pure-state controllability the system algebra has to be isomorphic to . However, the central result of this section shows that this is in general not necessary: a quasifree fermionic system (with or ) is fully pure-state controllable iff its system algebra is isomorphic to or , see Theorem 23.
6.1 Quasifree states
Due to the Wick expansion property, a quasifree state is uniquely characterized by its covariance matrix. (General references for this section include [20, 57–60].) The following proposition resumes a known result on these covariance matrices (see, e.g., Lemma 2.1 and Theorem 2.3 in ), which will be useful in the later development:
Proposition 17 The singular values of the covariance matrix of a d-mode fermionic state must lie between 0 and 1. Conversely, for any skew-symmetric matrix with singular values between 0 and 1 there exists a quasifree state that has as a covariance matrix.
6.2 Orbits and stabilizers of quasifree states under the action of quadratic Hamiltonians
The action of the time-evolution unitaries generated by quadratic Hamiltonians on quasifree states can be described by the next proposition (see Lemma 2.6 in ):
is defined using the skew-symmetric matrix T and generates the time-evolution of . The time-evolved state (at unit time), is again a quasifree state with a (skew-symmetric) covariance matrix , where .
using a (not necessarily unique) element where denotes the singular values of G. This means that a quasifree state can be reached from another one by the action of quadratic Hamiltonians if their covariance matrices share the same singular values (including multiplicities). Let us now recall another result related to the singular values of the covariance matrices of pure quasifree states (Theorem 6.2 in , and Lemma 1 in ):
Proposition 19 A quasifree state ρ is pure iff the following (equivalent) conditions hold for its covariance matrix : (a) The rows (and columns) of are real unit vectors which are pairwise orthogonal to each other. (b) The singular values of are all 1.
Applying this result together with Proposition 18, we obtain the next theorem:
Theorem 20 The set of quadratic Hamiltonians acts transitively on pure quasifree states, and the corresponding stabilizer algebras are isomorphic to .
Proof We have already shown that the singular values of the covariance matrices (with multiplicities) form a separating set of invariants for the orbits generated by quadratic Hamiltonians over the set of quasifree states. This means, according to Proposition 19, that the pure quasifree states form a single orbit.
By noting that Ω and (with ) are linearly independent vectors, we can conclude that a quadratic Hamiltonian H leaves the Fock vacuum invariant iff . In Theorem 43 of Section 9 we will show that these operators form a Lie algebra isomorphic to . □
Corollary 21 Theorem 20 identifies the space of pure quasifree states with the quotient space .
6.3 Conditions for quasifree pure-state controllability
According to Theorem 20, a set of quasifree control Hamiltonians allows for quasifree pure-state controllability, if the corresponding Hamiltonians generate the full quasifree system algebra, i.e. if . It is natural to ask whether this condition is also a necessary. Remarkably, it turns out that this is not the case, which is shown by the following lemma:
Lemma 22 Consider a quasifree fermionic system with modes. Let K be the subgroup of which is isomorphic to and stabilizes the first coordinate; its Lie algebra is denoted by . Then (a) the group K acts via its adjoint action transitively on the set of all skew-symmetric covariance matrices of pure quasifree states (whose singular values are all 1) and (b) the quasifree system is pure-state controllable if its system algebra is conjugate under to .
where is a dimensional unit real vector and is a skew-symmetric matrix. Now the proof of (a) follows using the induction hypothesis. The statement (b) is a consequence of (a). □
We relate Lemma 22 to what is known about transitive actions on the coset space . Only Lie groups isomorphic to and can act transitively (i.e. in a pure-state controllable manner) on the homogeneous space assuming . The case is discussed in . For we have and (where denotes the complex projective space in four dimensions), and it is known that only subgroups of isomorphic to or can act transitively on (see p.168 of  or p.68 of ; refer also to ).
In most cases the -subalgebras of are conjugate to each other. More precisely, Lemma 7 of  states that for all subalgebras of whose dimension is equal to are conjugate to each other under the action of the group . In particular, it follows in these cases that all subalgebras of with dimension are isomorphic to . Interestingly, the last statement holds also for (see Lemma 3 of ); however not all of these subalgebras of are conjugate. We obtain the following theorem providing a necessary and sufficient condition for full quasifree pure-state controllability in the case of or modes:
Theorem 23 A quasifree fermionic system with or modes is fully pure-state controllable iff its system algebra is isomorphic to or .
Proof “⇒”: Note that Theorem 20 identifies the space of pure quasifree states with the homogeneous space . Assuming , we summarized above that a group acting transitively on this homogeneous space is isomorphic either to or . Thus only the full quasifree system algebra or a system algebra isomorphic to can generate a transitive action on the space of pure quasifree states.
“⇐”: As discussed, all -subalgebras are conjugate to each other for and . Lemma 22(b) then implies that any set of Hamiltonians generating a system algebra isomorphic to will allow for full quasifree pure-state controllability. □
Note that the cases and are well-known pathological exceptions. The algebra breaks up into a direct sum of two -algebras which hence cannot be conjugate to each other. For , there are three classes of non-conjugate subalgebras of type in where two classes are given by irreducible embeddings and the third one is conjugate to the reducible standard embedding fixing the first coordinate. (For details, refer to the discussions on the pp.57-58 of , on the pp.234-235 of , or on the pp.418-419 of . In addition, this information can also be inferred from the tables on p.260 of .)
On a more general level, Theorem 23 can be seen as a fermionic variant of the pure-state controllability criterion for spin systems [27–29]. We note here that the result for spin systems has been recently generalized from the transitivity over a set of one-dimensional projections (i.e. pure states) to the transitivity over a set of projections of arbitrary fixed rank (i.e., over Grassmannian spaces) . We will use exactly this generalization in Section 9.3 in order to find a necessary and sufficient pure-state controllability condition for particle-conserving quasifree systems.
7 Translation-invariant systems
System algebras of translation-invariant fermionic systems with d modes for (a) nearest-neighbor interactions only and (b) arbitrary translation-invariant interactions
System algebra for (a)
System algebra for (b)
In this context, two sets of natural questions arise: (a) How does the dimension of these dynamic system algebras scale with the number of modes? (b) How do the system algebras generated by the nearest-neighbor terms differ from the general translation-invariant ones? Can one characterize those elements that are translation-invariant yet not generated by nearest-neighbor Hamiltonians? Are there, for example, next-nearest-neighbor interactions of this type? In this section, we will answer these questions partially. We determine the system algebra for general translation-invariant fermionic Hamiltonians, and conclude that its dimension scales exponentially with the number of modes. We also provide translation-invariant fermionic Hamiltonians of bounded interaction length which cannot be generated by nearest-neighbor ones.
The structure of this section is the following: As the structure of system algebras for translation-invariant systems has only been studied sparsely even for simple scenarios of spin models, we start by examining this case first. In Sections 7.1 and 7.2, we determine the system algebras of all translation invariant spin-chain Hamiltonians with L qubits. In particular, we simplify and generalize results of  concerning finite-ranged interactions. Finally, we present the corresponding results for the fermionic case in Sections 7.3 and 7.4.
7.1 Translation-invariant spin chains
where . We will determine the translation-invariant system algebra which is defined as the maximal Lie algebra of skew-Hermitian matrices commuting with the translation unitary .
where denotes the greatest common divisor of L and k.
It follows that the number of -invariant basis vectors and—hence—the trace of is equal to . Thus, the multiplicities of are given accordingly by . □
Note that a Hamiltonian commutes with iff it commutes with all spectral projections of . Combining this fact with Theorem 51 we obtain a characterization of the system algebra for translation-invariant spin systems:
Theorem 25 The translation-invariant Hamiltonians acting on a L-qubit system generate the system algebra , where the numbers are defined in Eq. (30).
In complete analogy one can show that for a chain consisting of L systems with N levels, the system algebra is equal to , where denotes the Fourier transform of the function .
7.2 Short-ranged spin-chain Hamiltonians
In many physical scenarios, we may only have direct control over translation-invariant Hamiltonians of limited interaction range. We will investigate in this section how the limitations on the interaction range constrain the set of reachable operations. In particular, we provide upper bounds for the system algebras with finite interaction range.
for all combinations of apart from the case when . In this way, corresponds to the translation-invariant on-site Hamiltonians, while is generated by the on-site terms and the nearest-neighbor interactions, and so on. Finally, we have .
System algebras of translation-invariant systems with spins and interaction lengths of less than M . Refer also to Theorem 25 for the structure of
L = 1
L = 5
We first mention a central proposition whose proof can be found in Appendix E:
Proposition 26 Let denote a divisor of L. Given two elements and , we obtain that and hold for any positive integer q.
Applying Proposition 26, we can present upper bounds for the system algebras with restricted interaction length.
Theorem 27 Let denote a divisor of the number of spins L, and define . We obtain: (a) The algebra is isomorphic to a Lie subalgebra of and does not generate . (b) The algebra is isomorphic to a Lie subalgebra of and does not generate . (c) In addition, .
Obviously, holds. Using the formula for in Appendix E.2, we obtain that holds for every integer q. Hence, . □
In particular, this theorem implies that the algebra of all translation-invariant Hamiltonians cannot be generated from the subclass of nearest-neighbor Hamiltonians, cf. also . More precisely, one finds:
Corollary 28 If L is even, is isomorphic to a Lie subalgebra of the Lie algebra . For odd , is isomorphic to a Lie subalgebra of the Lie algebra .
Let us now compare our upper bounds with the results of Table 2. Theorem 27 restricts the possibilities for the M-local algebras only by some central elements when compared to the corresponding full translation-invariant algebra . One can indeed identify in Table 2 some missing -parts for . In general, the dimensions of the M-local algebras can be even smaller than predicted by the upper bounds of Theorem 27 as can be seen in Table 2 for . Theorem 27 and Table 2 suggest that the prime decomposition of the chain length L will have strong implications on the dimension of .
7.3 Translation-invariant fermionic systems
The superselection rule for fermions splits the spectral decomposition of the translation unitary into two blocks corresponding to the positive and negative parity subspace. The translation unitary commutes with the parity operator P, and hence is block-diagonal in the eigenbasis of P where and . The following lemma gives the spectral decomposition of the operators :
If is even, the sum is even for all of the bit strings invariant under a cyclic shift by k positions. It follows that all the diagonal entries of are zero, while has non-zero diagonal entries. The non-zero diagonal entries of are given by the number s of Eq. (36). Note that s is +1 if is even; and −1 otherwise. Hence the frequencies of +1 and −1 in the set of diagonal entries are equal. In summary, if is even.
Assume now that is odd. The sum is odd for half of the bit strings and even for the other half. Applying again Eq. (36), we obtain always a positive sign. Hence, both traces are equal to . This completes the proof. □
Lemma 29 together with Theorem 51 implies the following characterization of the system algebra for a translation-invariant fermionic system:
where the numbers are defined in Eq. (35).
Remark 31 Note that holds for any ℓ and that . It follows that and hence that the dimension of the system algebra in Theorem 30 scales exponentially with d.
where is guaranteed to be an integer by Fermat’s little theorem.
7.4 Fermionic nearest-neighbor Hamiltonians
For spin systems (see Section 7.2) we verified that the translation-invariant nearest-neighbor interactions together with the on-site elements will never generate all translation-invariant operators, i.e. (if the number of spins L is greater than two). This means that there exist certain translation-invariant elements which cannot be generated by nearest-neighbor interactions and on-site elements, but we could not identify the explicit form of these translation-invariant elements for general L. In particular, it would be interesting to know if holds for interaction lengths less than M (), where M is independent of L.
In the case of fermionic systems, we can provide a result in this direction due to the restriction imposed by the parity superselection rule, which strongly limits the set of nearest-neighbor Hamiltonians. As we have discussed at the beginning of this section, the fermionic translation-invariant Hamiltonians of nearest-neighbor type are spanned by only six elements: , , , , , and as defined in Eqs. (25)-(28). We can show that there exist next-nearest-neighbor or third-neighbor interactions for odd which cannot be generated by these six Hamiltonians, while for even we provide a fourth-neighbor element.
Let denote the subalgebra of (see Theorem 30) which is generated by all elements of interaction length less than M. In particular, is generated by nearest-neighbor and on-site elements. The result of this subsection is summarized in the following theorem:
The generator is not contained in the system algebra generated by nearest-neighbor interactions and on-site elements if is odd, while the element is not contained in if is even. Hence (when ).
Note that the Hamiltonian of Theorem 33 is a third-neighbor Hamiltonian for and a next-nearest-neighbor Hamiltonian for . The proof of Theorem 33 is rather involved. The proof for even d is given in Appendix F, while Appendix G contains the proof for odd d.
8 Quasifree fermionic systems satisfying translation-invariance
System algebras of quasifree fermionic systems with d modes satisfying translation-invariance
General case (see Theorem 34)
(Twisted) Reflection symmetry (see Eq. (45) and Corollary 35)
2n − 1
8.1 Translation-invariant quadratic Hamiltonians
Note that the operators (for d even), , , , , and are linearly independent and span the -dimensional space of all translation-invariant quadratic Hamiltonians. For notational convenience we also introduce the dummy operators (assuming d is even) and for .
With these stipulations, we can characterize the system algebra:
Theorem 34 Let denote the system algebra on a fermionic system with d modes which corresponds to the set of Hamiltonians that are translation-invariant and quadratic. Then the Lie algebra is isomorphic to for odd d and to for even d.
provides an explicit Lie isomorphism between and . If is even, the system algebra consists of the above-described generators supplemented with the element . This additional element commutes with all the other generators and—therefore—provides an additional . □
A given translation-invariant quasifree Hamiltonian is ℛ-symmetric (i.e. ) iff the coefficient matrix is restricted to be real. In our language, these Hamiltonians are exactly the ones for which , i.e., the corresponding generators are spanned by the operators (for d even), , , , and . From the proof of Theorem 34 one can immediately deduce the corresponding system algebra:
Corollary 35 Consider a fermionic system with d modes and the set of quadratic Hamiltonians which are translation-invariant and ℛ-symmetric. The corresponding system algebra is isomorphic to for odd d and to for even d.