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A dynamic systems approach to fermions and their relation to spins
EPJ Quantum Technologyvolume 1, Article number: 11 (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
As a starting point, consider the controlled Schrödinger (or Liouville) equation
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 .
For a N-level system, the natural representation as Hermitian operators over relates the Hamiltonians as generators of unitary time evolutions to the Lie algebra of skew-Hermitian operators that generate the unitary group of propagators. Let be a subset of Hamiltonians seen as Lie-algebra elements. Then the smallest subalgebra (with respect to the commutator ) of containing L is called the Lie closure of L written as . Moreover, for any element , there exist control amplitudes with such that (and similarly with a drift term)
where denotes time-ordering.
Now taking the Lie closure over the system Hamiltonian and all control Hamiltonians of a bilinear control system defines the dynamic system Lie algebra (or system algebra for short)
It is the key to characterize the differential geometry of a dynamic system in terms of its complete set of Hamiltonian directions forming the tangent space to the time evolutions. For instance, the condition for full controllability of bilinear systems can readily be adopted from classical systems [31–34] to the quantum realm such as to take the form of
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 .
Remark 1 Interestingly, this notion is stronger than the requirement that from any given (normalized) pure state one can reach any other (normalized) pure state, since it is well known [27–29] that for N being even, all rank-one projectors are already on the unitary symplectic orbit
and is a proper subgroup of .
In general, the reachable set to an initial state of a dynamic system with system algebra is given by the orbit of the dynamic (sub)group generated by the system algebra
Thus the system algebra can be envisaged as the fingerprint encoding all the dynamic properties of a dynamic system Σ. Via the respective reachable sets (see, e.g., ) it is easy to see that a coherently controlled dynamic system can simulate the dynamics of another system if and only if the system algebra of the simulating system encompasses the system algebra of the simulated system ,
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.
Next we describe dynamic symmetries of bilinear control systems whose Hamiltonians are given by . The symmetry operators s are collected in the centralizer of in :
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.
Likewise the symmetries to a given set of states are given by its centralizer
where denotes the real span. Clearly, generates the stabilizer group to the state space of the control system .
Since in the absence of other symmetries the identity is the only and trivial symmetry of both any state space as well as any set of Hamiltonians and their respective system algebra , one has . So there is always a trivial stabilizer group . This explains why the time evolutions generated by two Hamiltonians and coincide for the set of all density operators if (and without other symmetries only if) . As is well known, by the same argument, in time evolutions
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:
Lemma 2 Consider a bilinear control system with system algebra on a state space . Let and while assuming that for all , i.e., operations generated by map the set into itself. Then the condition
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
The complex Hilbert space of a d-mode fermionic system with one-particle subspace is the Fock space
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 ().
The fermionic creation and annihilation operators, and act on the Fock space in the following way: , , , and ; while in the general case of , their action is given by and . By their definition, these operators satisfy the fermionic canonical anticommutation relations
where denotes the anticommutator. Moreover, every linear operator acting on can be written as a complex polynomial in the creation and annihilation operators.
Another set of polynomial generators acting on the Fock space is given by the 2d Hermitian Majorana operators and , which satisfy the relations ()
A product of Majorana operators is called a Majorana monomial. The ordered Majorana monomials with form a linearly independent basis of the complex operators acting on . Each Majorana monomial acting on d-mode fermionic system can be identified with a complex operator acting on a chain of d qubits via the Jordan-Wigner transformation [35–38] which is induced by
where the following notation for the Pauli matrices , , and is used.
Now we highlight the real subspace contained in the set of complex operators acting on the Fock space which consists of all skew-Hermitian operators and which forms the real Lie algebra closed under the commutator and real-linear combinations. More precisely, is generated by all operators
where M denotes any ordered Majorana monomial and
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.
The parity superselection rule identifies among the operators acting on the physical observables as those that do commute with the parity operator
where the adjoint action of P on a Majorana monomial is given as
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.
Now we will discuss why the set of all physical fermionic states consists similarly of all density operators that commute with P, notably . As we will show, the parity superselection rule induces a decomposition into a direct sum of two irreducible state-space components exploiting . Recall that and the eigenspaces to the eigenvalues +1 and −1 are indeed of equal dimension, as there are exactly even operators which map the vacuum state Ω into the +1 eigenspace of P. Note that . Thus the Fock space can be split up as a direct sum of two equal-dimensional eigenspaces of P, called the positive and negative parity subspaces:
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.
Now we may write with the orthogonal projections and projecting onto the respective subspaces. Any physical observable (i.e. even operator) A has a block-diagonal structure with respect to the above splitting, i.e. . This follows, as the requirement enforces for any operator . We obtain
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
Here we derive a general controllability result for fermions obeying the parity superselection rule. We illustrate that full controllability for a fermionic system can be achieved with quadratic Hamiltonians and a single fourth-order interaction term. For example, in a system with d modes, the complete fermionic dynamical algebra (see Theorem 4) can be generated by a quartic interaction between the first two modes combined with three quadratic Hamiltonians: the nearest-neighbor hopping term
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 .
This also implies that for any physical fermionic Hamiltonian H, there exists a unique Hamiltonian
that is traceless on both the positive and the negative parity subspaces, i.e.,
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):
Proposition 6 Consider a fermionic quantum system with modes. The system algebra of a fully controllable fermionic system can be generated using the operators , , , and with the map L as defined in Eqs. (12) and (13), where
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 guide the discussion in a different direction by emphasizing that the property does not determine the system algebra uniquely. We define the centralizer of a set in (e.g. ) as
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:
Corollary 8 Let denote an irreducible subalgebra of , i.e. . Then one finds that if and only if .
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.
In the Majorana monomial basis, the quadratic Hamiltonian H can be written as
The properties of A and B directly imply that the matrix T is real and skew-symmetric. Using the formula
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
We start by showing that the full system algebra of quasifree fermions can be generated using only three quadratic operators, namely , , and from Eqs. (17a) and (17b) where , , and . The Jordan-Wigner transformation maps these generators respectively to the Heisenberg-XX term
, and , where operators as (e.g.) are defined as .
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.
For illustration, note that two elements in the commutant are trivial, to wit the identity and the generator for the swap-operation between the two tensor copies. The third element does not yet occur in the unitary case described in : it is the projector onto the totally anti-symmetric state. To see this, recall that  implies that if the Hamiltonians generate a system algebra of orthogonal type, then there is an operator satisfying
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
A fermionic state ρ on is called quasifree or Gaussian if it vanishes on odd monomials of Majorana operators and factorizes on even monomials into the Wick expansion form
Here the sum runs over all pairings of , i.e., over all permutations π of satisfying and for all q. The covariance matrix of ρ is defined as the skew-symmetric matrix with real entries
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 ):
Proposition 18 Consider a quasifree state corresponding to the (skew-symmetric) covariance matrix . The quadratic Hamiltonian
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 .
Any skew-symmetric matrix G can be brought into its canonical form
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.
As the set of quadratic Hamiltonians generate a transitive action over the pure quasifree states, the corresponding stabilizer subalgebras are isomorphic to each other. Consider a quadratic Hamiltonian H with the coefficient matrices A and B as given in Eq. (18) and the Fock state , which is the projection onto the Fock vacuum vector Ω. The state is left invariant under the time evolution generated by H () iff Ω is an eigenvector of H. We obtain that
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 .
Proof We prove (a) by showing that all pure quasifree states can be transformed under K-conjugation to the same pure state. We employ an induction on d. The base case can be directly verified. It follows from Proposition 19(b) that the skew-symmetric covariance matrix of a pure quasifree state can be written as , where denotes a normalized -dimensional vector and denotes a -dimensional skew-symmetric matrix. We consider the action of a general orthogonal transformation with :
Since any -dimensional vector with unit length can be transformed by an orthogonal transformation to , we can choose such that . We have as the transformed matrix is skew-symmetric. Again by Proposition 19(b) we obtain the transformed matrix as
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
We study system algebras generated by translation-invariant Hamiltonians of the type which arises approximately in experimental settings of, e.g., optical lattices. As the naturally occurring interactions are usually short-ranged, we pay particular attention to the case of Hamiltonians with restricted interaction length. For example, consider a d-site fermionic chain with Hamiltonians which are translation-invariant and are composed of nearest-neighbor (plus on-site) terms. All elements in its dynamic algebra can be written as linear combinations of six types of terms: the chemical potential
the real and complex hopping Hamiltonians
the real and complex pairing terms
as well as a local density-density-type interaction
The corresponding dynamic system algebras (given in Table 1) were computed with the help of the computer algebra system magma  for up to six modes while distinguishing nearest-neighbor interactions from arbitrary translation-invariant ones.
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
Consider a chain of L qubits with Hilbert space . The translation unitary is defined by its action on the canonical basis vectors as
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 .
Lemma 24 The translation unitary can be spectrally decomposed as , and the rank of the spectral projection is given by the Fourier transform
where denotes the greatest common divisor of L and k.
Proof The eigenvalues of are limited to with as the order of is L, i.e. . Hence, the corresponding spectral decomposition is given by . This induces a unitary representation of the cyclic group which maps the k th power of the generator of degree L to . Note that splits up into a direct sum containing copies of the one-dimensional representations satisfying . Therefore, we determine the rank of a projection by computing the multiplicity of using the character scalar product
The trace of is equal to the number of basis vectors left invariant since is a permutation matrix in the canonical basis. From elementary combinatorial theory we know that a bit string is left invariant under a cyclic shift by k positions if and only if it is of the form
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.
Let us denote the Lie algebra corresponding to Hamiltonians of interaction length less than M by , or for short. In other words, is the Lie subalgebra of generated by the skew-Hermitian operators
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 .
We computed all the algebras for and using the computer algebra system magma . The results, shown in Table 2, suggest that for certain restrictions on the interaction length (e.g., nearest-neighbor terms), there will be some translation-invariant interactions that cannot be generated. This is in accordance with the result of Kraus et al. . Building partly on their work, we analyze the properties of the algebras for general M and L values, and then compare our theorems with Table 2.
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, .
Proof (a) Since M is a divisor of L, the equation
holds for any integer q, and one can invert the equation as . If , we obtain by applying Proposition 26 that
holds for . It follows that is a subalgebra of the Lie algebra which consists of all skew-Hermitian matrices satisfying the condition in Eq. (31). Note that is isomorphic to , and part (a) follows. (b) For elements , Proposition 26 and Eq. (31) imply that
The maximal Lie algebra consisting of skew-Hermitian matrices which satisfy the condition in Eq. (32) is isomorphic to . (c) Let
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
To determine the system algebra generated by all translation-invariant Hamiltonians of a fermionic chain, we can follow similar lines as in Section 7.1. Here, however, we additionally have to consider the parity superselection rule. We define the fermionic translation-invariant system algebra as the maximal Lie subalgebra of [see Theorem 4] which contains only skew-Hermitian matrices commuting with the fermionic translation unitary , which is defined below such that it commutes with the parity operator P (see Eq. (14)). The standard orthonormal basis in the Fock space for a chain of d fermionic modes is given by
with . Note that for the purpose of unambiguously defining this basis, we order the operators in Eq. (33) with respect to their site index i. The fermionic translation unitary is defined by its action
on the standard basis. The adjoint action of on the creation operators is then given by
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 :
Lemma 29 The unitary operators can be spectrally decomposed as , where the rank of the spectral projection is given by the Fourier transform
of where and
Proof We determine the spectral decomposition of and along the lines of Lemma 24. Let denote the subspace spanned by those basis vectors of Eq. (33) for which is even. Likewise, corresponds to the case of odd . As , the eigenvalues of are of the form with . Hence, the spectral decomposition is given by . We define representations of the cyclic group which map the k th power of the generator of degree d to . Note that splits up into a direct sum containing copies of the one-dimensional representations satisfying . The rank of the projection is equal to the multiplicity of in the decomposition of the reducible representation . This multiplicity can be computed as the character scalar product
In the standard basis, all matrix entries of are elements of the set . It follows by repeated applications of Eq. (34) that maps the basis vectors to where π is a cyclic shift by k positions and the sign s is given by
Recall from the proof of Lemma 24 that a bit string is left invariant under a cyclic shift by k positions iff it is of the form
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:
Theorem 30 Let the translation-invariant Hamiltonians act on a fermionic system with d modes. The corresponding system algebra is given by
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.
Remark 32 Assuming that the number of modes is given by a prime number p, we can explicitly determine the numbers from Eq. (35). The corresponding system algebras are
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:
Theorem 33 Let us consider the Hamiltonian , and fourth-neighbor Hamiltonian
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
We continue the discussion of translation-invariant fermionic systems from Section 7 by narrowing the scope to quadratic Hamiltonians. In Section 8.1, we derive the dynamic algebras for systems with and without (twisted) reflection symmetry. Both of these cases are summarized for quasifree fermionic systems in Table 3: the system algebras were computed using the computer algebra system magma  for cases with low number of modes, while the complete picture is provided by Theorem 34 and Corollary 35. Section 8.2 yields a classification of the orbit structure of pure translation-invariant quasifree states. This allows us to present an application to many-body physics in Section 8.3, where we bound the scaling of the gap for a class of quadratic Hamiltonians.
8.1 Translation-invariant quadratic Hamiltonians
A quadratic Hamiltonian H is translation-invariant (i.e. ) iff the coefficient matrices A and B in Eq. (18) are cyclic (i.e. ). To study such Hamiltonians, it is useful to rewrite them in terms of the Fourier-transformed annihilation and creation operators
with , which satisfy the canonical anticommutation relations
A Hamiltonian from Eq. (18) with cyclic A and B can now be rewritten as
applying and , as well as the notation . The hermiticity of A and the skew-symmetry of B translates into the properties and . This allows us to decompose the Hamiltonian into a four-part sum
where one has the following definitions
with as well as
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.
Proof If is odd, the generators , , , , and can be partitioned into m pairwise-commuting sets, which each span linear subspaces as
with . The commutation properties (with ) follow from Eq. (39). Moreover, is one-dimensional and forms a -algebra. Using Eq. (39), the relations and can be deduced for . Substituting k and into a and b in the previous formula, one can verify directly that the correspondence
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 . □
The isomorphism between and as given in the proof leads to a compact formula for the time evolution (in the Heisenberg picture) of the elements of . Since the operators , , , and (with ) satisfy the same commutation relations as the Pauli matrices X, Y, Z, and , their time-evolution generated by the Hamiltonian H in Eq. (41) can be straightforwardly related to a qubit time-evolution
The twisted reflection symmetry plays an important role in translation-invariant quasifree fermionic systems. It is defined by the unitary
whose adjoint action on creation operators and their Fourier transforms is given by
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.
Given the system algebras and , we investigate the subalgebras generated by short-range Hamiltonians. It will be useful to introduce for the Hamiltonians
as well as the additional ones ( only for even d)
In these definition we used cyclic indices, e.g. . The operators (for d even), , , , , and span linearly. Using the identities above, the commutation relations of the operators, and some trigonometric identities, we obtain
for . In  it was shown that already the nearest-neighbor Hamiltonians of generate the whole . Now we are in the position to provide a more systematic proof of their result:
Lemma 36 The system algebra can be generated using the one-site-local operator and a nearest-neighbor element with assuming that or for odd d and additionally requiring for even d.
Proof (1) From Eqs. (48a)–(48d) we know that , , , and would generate the whole . (2) Suppose that and . From and it follows that one can generate , , , and . Hence according to observation (1), the whole is generated. (3) Suppose now that , d is odd, and . From one can generate , , and . From Eqs. (48a)–(48d) it follows that these generators in turn generate all . Since d is odd, is also generated. Hence we obtain , , , and , and according to (1), the algebra is generated. □
For the more general , we obtain a slightly larger system algebra when we do not assume ℛ-symmetry:
Proposition 37 The elements of with interaction length less than M (where and ) generate a system algebra which is isomorphic to for odd d and to for even d.
Proof From Lemma 36 we know that the operators with already generate which is isomorphic to for odd d and to for even d. We have additional operators with which are linearly independent and commuting. These generate the other parts corresponding . □
We illustrate Lemma 36 and Proposition 37 with a fermionic ring of modes. Suppose that the drift Hamiltonian of this system is the nearest-neighbor hopping Hamiltonian , and that one can additionally control the on-site potential , the pairing strength , and the magnetic flux in the ring. Lemma 36 implies that the first three Hamiltonians generate the Lie algebra of all Hamiltonians which are simultaneously ℛ-invariant, translation-invariant, and quadratic. The magnetic flux term commutes with all elements of and contributes only an additional to the system algebra. Thus, the system algebra generated by all nearest-neighbor quadratic Hamiltonians that are translation-invariant is given by . In order to achieve full controllability for a translation-invariant and quasifree fermionic system (which corresponds to the Lie algebra ), one has to add a next-nearest neighbor Hamiltonian as .
8.2 Orbits of pure translation-invariant quasifree states
We characterize now the orbits of pure translation-invariant quasifree states under the action of translation-invariant quadratic Hamiltonians. Since the operators commute with all the other translation-invariant quadratic Hamiltonians (as discussed in Section 8.1), their expectation values stay invariant under the considered time evolutions. At the end of the section, we show that these invariant expectation values even form a separating set of invariants for the orbits of pure translation-invariant quasifree states.
Let us recall that a quasifree state is fully characterized by its Majorana covariance matrix, defined in Eq. (24). The translation unitary acts on the Majorana operators by conjugation as . It follows that a quasifree state ρ is translation-invariant (i.e. ) iff its covariance matrix is doubly-cyclic, i.e. . The double-cyclicity of G implies that it can be expressed as a block-Fourier transform of a block-diagonal matrix, i.e.
where with and with being -matrices. The matrices can be calculated by the inverse block-Fourier transform
The fact that G is skew-symmetric and real implies
Moreover, due to Eq. (49) the set of eigenvalues of all the matrices equals the one of (including multiplicities). Combining these observations with Proposition 17 and Proposition 19, we obtain the following characterization of pure translation-invariant quasifree states:
Lemma 38 A set of matrices (with ) defines a covariance matrix of a pure quasifree state through Eq. (49) iff they satisfy Eq. (51) and their eigenvalues are in the set .
The entries of and the expectation values of the operators defined in Eq. (42) can be related by
using Eq. (50) and the definitions for , , , and . Now we can prove the main theorem of this subsection:
Theorem 39 Two pure quasifree states and can be connected through the action of a translation-invariant quadratic Hamiltonian if and only if holds for all with .
Proof First, we consider the ‘if’-case: Let H be a translation-invariant quadratic Hamiltonians for which holds. Since the operators commute with any translation-invariant Hamiltonian, we have . Second, we treat the ‘only if’-case: Let and denote the Fourier-transformed Majorana two-point functions (defined as in Eq. (50)) of and , respectively. The action of a translation-invariant Hamiltonian, is represented by the map
where is given by . Using Eq. (52), we obtain for . These expectation values have to be in the set , since the eigenvalues of and are in the set . Then, it follows from that the expectation values of and coincide. Thus, we obtain from Eq. (53) that and can be transformed into each other. □
Finally, we turn to the ℛ-symmetric setting, as introduced in Section 8.1, and determine the orbit structure of quasifree pure states which are translation-invariant and ℛ-symmetric under the action of operators in .
Proposition 40 The unitaries generated by the Lie algebra act transitively on the set of quasifree pure states which are translation-invariant and ℛ-symmetric.
Proof Since , the expectation value of these operators in ℛ-symmetric states must vanish as . Moreover, by Theorem 39 we know that two pure translation-invariant states are on the same -orbit iff the expectation values of the operators coincide for all . Hence the translation-invariant ℛ-symmetric states lie on the same -orbit. As Eq. (53) implies that the -orbits are equivalent to -orbits, we have proved the proposition. □
8.3 An application to many-body physics
In many-body physics, one of the important characteristics of quantum criticality is the closing of the gap. This means that the energy difference between the ground state and the first excited state goes to zero in the thermodynamic limit, when the number of spins or fermionic modes goes to infinity. Quasifree fermionic models can display both gapped and gapless behavior. Using the techniques developed in the previous subsections, we will prove that the gap always disappears (i.e. closes) for translation-invariant quasifree models if the coefficient matrix A of Eq. (18) is purely imaginary while B is an arbitrary, complex skew-symmetric matrix. Different cases have been considered in .
To formalize this statement, let us consider a set of fixed (finite) real numbers with and a set of fixed complex numbers (of finite modulus) with . With these stipulations, we define for any the cyclic matrices and (or A and B for short) by specifying their entries
By applying these definitions to Eq. (18) we obtain:
Theorem 41 Given the positive integers d and M with , consider the corresponding translation-invariant quasifree Hamiltonian
where A and B are defined in Eqs. (54) and (55). Assume that has a unique ground state. Then the gap of is bounded by , i.e. the gap closes algebraically in the thermodynamic limit of d going to infinity.
Proof Since is translation-invariant and its coefficient matrix is imaginary, it can be decomposed in terms of the operators with and as
using , , as well as . Let be a pure quasifree state, and let denote its Fourier-transformed Majorana two-point functions (see Eq. (50)). From Eq. (44) we know that is an eigenstate of iff . The eigenvalue of corresponding to this state is given by
Let us emphasize that the proof builds on the fact that M is fixed and finite, while d goes to infinity in the thermodynamic limit. Among the eigenstates of , consider the (unique) ground state , whose Fourier-transformed Majorana two-point functions (see Eq. (50)) will be denoted by . From this ground state let us construct another quasifree state which is defined through its Majorana two-point functions
while for general we assign .
The corresponding pure quasifree state is an eigenstate of , since according to Eq. (53) its Fourier-transformed Majorana two-point function stays invariant during the time-evolution generated by . Using Eq. (56), we can calculate the difference between the energies corresponding to and as