is exact, dimF2kerδ= dimF2imδ0. As we have seen in Example 4.2,
Tor1(T, B)∼= Tor1(R/m, B)∼= (F2)3, and Tor1(N, B)∼= Tor2(R/m, B)∼= (F2)3. It follows that dimF2K(L)≤dimF2Tor1(N, B) + dimF2Tor1(T, B) = 6.
It is routine to verify that b4 ⊆I2(φ) ⊆m := (x+ 1, y+ 1, z+ 1). Recall the decomposition K(L) = L
pK(L)p where p runs over all maximal ideals of R/bL. Due to Lemma 3.4.1, this decomposition consists of only one summandK(L)m. WhenLis odd, since (bL)m =mm, we know K(L)m=K(1)m. Sinceφ7→0 under a=b=c= 0, we see dimF2K(1) = 4. The logical operators in this case are
0 0 zb zb
^xb·ycz¯
0 1 0 0
;
xb xb 0 0
^bz·xyc
1 1 1 0
where µb=PL−1
n=0µn so µ·bµ=bµ, and symplectic pairs are tied. The left elements are string-like, and the right surface-like.
WhenLis even, the following areF2-independent elements ofK(L). As there are 6 in total, the largest possible number, we conclude thatK(L) is 6-dimensional, i.e., the number of encoded qubits is 3 when linear dimensions are even.
0 0 bz bz
^bx0yb0
1 +y x+xy
0 1 +x+y+xy
;
xb xb 0 0
^yb0zb0
1 +z 1 +z 1 +z y+yz
;
yb yb yb yb
^bz0xb0
0 1 +x+z+xz
1 +x 1 +x
where bµ0 =PL/2−1
i=0 µ2i so (1 +µ)µb0 =bµ. The pairs are transformed cyclically byx7→y 7→z 7→x
together with the symplectic transformationω of Eq. (4.3). ♦
It is interesting on its own to prove or disprove that the elementary symplectic transformations generate the whole symplectic transformation group. In the zero-dimensional case where Λ is the trivial group, it is true as we have already seen in Proposition 2.1. The one-dimensional case is also true because it is implied by the computation in the proof of Theorem 4.1. A classical problem answered affirmatively by Suslin [91] is that any sufficiently large invertible matrix over a polynomial ring is a finite product of elementary matrices such as row operations and scalar multiplications. Later, an algorithmic proof is given by Park and Woodburn [92]. A similar problem under a confusingly similar name ‘symplectic group’ over polynomial rings is solved by Grunewaldet al.[93], who defined the ‘symplectic group’ as
S ∈Mat (n,F[x1, . . . , xn])
STλS=λ whereT is the transpose. Kopeyko [94] generalized it to include Laurent polynomials, but still the ‘symplectic group’ is different from ours since the antipode map is absent from the definition
S∈Mat n,F[x±11 , . . . , x±1n ]
STλS=λ .
The characteristic dimension is not proven to be invariant under coarse-graining. It suffices to have an upper bound on dimF2K(L) in Corollary3.4.8. without the condition that 2-L. A closely related object is the Hilbert function. Given a graded moduleM =L∞
s=0Msover a polynomial ring with coefficients in a field F, the Hilbert function fM is a numerical function defined by fM(s) = dimFMs. SinceK(L) = dimF2Tor1(coker, R/bL), the Hilbert function might be useful if we could make cokergraded. A technical difficulty would be that the idealbL= (xL1 −1, . . . , xLD−1) is not a power ofm= (x1−1, . . . , xD−1). See [76, Chapter 12] and [95].
Lastly, an important problem is to give a criterion to a moduleM that can be realized as cokerσ† for an exact code Hamiltonian described byσ. In the two-dimensional case, we know the answer — annM = (x−1, y−1) by Theorem 4.2.
Chapter 5
Cubic code
The toric code [3] is usually defined on a planar graph with qubits residing on edges. The star operator and plaquette operators are defined according to the data of the graph. The model is thus a priori lattice dependent. However, a certain set of important properties of the model turns out to be lattice independent. Particularly the ground-state subspace conveys a structure that depends only on the topology of the underlying space, insensitive to the microscopic detail. If the underlying space is a genusg (oriented) surface, the degeneracy is 4g. Even if the Hamiltonian is perturbed, the energy splitting of lowest 4g states is exponentially small in the system size as long as the perturbation is small enough [5,3, 6]. It is a signature oftopological order.
Authors have used the term ‘topological quantum order’ to mean all or a part of the following properties: Ground-state degeneracy as a function of topology, no spontaneous symmetry breaking, anyonic particle content [24,96,97,5,26,25], robust edge modes against perturbations [98,99,37], locally indistinguishable ground states [6, 75, 74], and topological entanglement entropy [27, 28].
Here, we take the local indistinguishability of ground states as a definition of topological quantum order. The local indistinguishability is precisely the one used in Chapter 3, as well as in the proof of gap stability results [75,74]. See Lemma3.3.1. We ask what quantum phases are possible in the class of code Hamiltonians.
An important example of topological quantum order presented in this thesis iscubic code. In this chapter we explain how the model is found, and study its consequences. The cubic code is an exact local additive code with translation symmetry on the simple cubic lattice, where the exactness is as in Chapter3. This model is a gapped unfrustrated spin Hamiltonian, and is topologically ordered in the very sense we just defined, but breaks many aspects of conventional models of topological order.
Most prominently, the excitations or charges of the cubic code cannot be interpreted as particles since they are immobile; the hopping term appears only after L-th or higher order perturbation theory, where L is the linear system size. The immobility implies that the system spends quite a long time to reach its thermal equilibrium in response to environment’s change. (Later, we will give quantitative statements regarding this.) Moreover, due to the topological quantum order, the
phase is stable with respect to arbitrary but small perturbations; the immobility of excitations is also protected [6].
The immobility of the charges translates into the theory of quantum error correcting codes as the absence of string logical operators. We call this property byno-strings rule. The notion of string logical operators might be intuitive if one imagines the toric codes in two or three dimensions [7].
However, this intuition is too model specific; the strings in discrete lattices are not well-defined objects since a set of points in the lattice does not in general have a well-defined dimensionality. We overcome this issue by definingstring segments that capture characteristics that are responsible for the mobility of charges.
We proceed by translating conditions for the absence of the logical string segments into our algebraic framework, in order to systematically search for codes without logical strings; the cubic code is not an ad hoc model. By Corollary4.3.3[47], the characteristic dimension of the cubic code must be 1 and the degeneracy must generally grow with the system size. An explicit formula for the degeneracy is given. Additionally, a real-space renormalization computation is presented. It seems that the cubic code is a fixed point of a certain unconventional kind. Namely, the model is a direct sum of two daughter models Aand B at a coarse-grained lattice where one daughter model A is the same as the original model, but the otherB is not. The model B produces two copies of itself at a further coarse-grained lattice. The renormalization group flow continues to branch. Next, we compute the thermal partition function, and show that there is no finite temperature phase transition.
5.1 String segments and no-strings rule
The most important property of one-dimensional objects is that a finite part of it has two disjoint boundary points. An intuitive role of string operator is to move an excitation at its one boundary point to another boundary point. Essentially, it is a concatenation or juxtaposition of hopping operators. The hopping operator is, as a whole, a finitely supported operator. Therefore, if the hopping operator acts on the vacuum (ground state), the overall effect is to create a trivial or neutral charge, which consists of two spatially separated charges. When either of the two charges is neutral by itself, the hopping is meaningless because trivial charges can be annihilated locally and created anywhere arbitrarily; only the hopping of nontrivial charges is important. Now we can define string segments and their absence.
Definition 5.1. [50, 51] A string segment is a finitely supported Pauli operator that creates excitations contained in the union of two finite boxes ofwidthw. The string segment isnontrivial if the charge contained in one of the boxes is nontrivial. The distance between the boxes is the lengthof the string segment. We say a model obeysno-strings ruleif the length of any nontrivial
string segment of widthwis bounded byαw for some constantα≥1.
The no-strings rule may seem too strong than necessary; why do we need an upper bound by alinear function? It is rather a technicality that is necessary to prove a logarithmic energy barrier theorem in Section6.1. However, our definition seems sharp yet broad enough to derive further results.
An immediate consequence of the no-strings rule in the case of a translation-invariant code Hamiltonian is that there are infinitely many charges. If there were only finitely many, then in the sequence of all translations of a chargec there would be an equivalent chargec0. It meansc−c0 is neutral and thereforec−c0 is created by a finitely supported operator, which is a nontrivial string segment. This string segment can be juxtaposed many times to give arbitrarily long string segments with fixed width. This is a contradiction to the no-strings rule. Also, in view of Corollary4.3.3and Lemma 3.4.9, a translationally invariant three-dimensional exact code Hamiltonian that obeys the no-strings rule, must have a growing ground-state degeneracy with respect to the system size when defined on lattices with periodic boundary conditions.