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Complete set of commuting observables

In quantum mechanics, a complete set of commuting observables (CSCO) is defined as a maximal collection of Hermitian operators that pairwise commute, such that their simultaneous common eigenvectors form a complete, orthonormal basis for the system's Hilbert space, allowing each basis state to be uniquely labeled by the set of corresponding eigenvalues.^[1]^[2]^[3] This concept arises from the compatibility theorem, which states that commuting observables can be simultaneously diagonalized in a shared eigenbasis, enabling precise, non-disturbing measurements of all members of the set on the same quantum state.^[1]^[2] The completeness condition ensures that no further commuting observable can resolve any remaining degeneracies in the eigenspaces, making the set maximal and sufficient to fully specify the quantum state without ambiguity.^[2]^[3] The importance of a CSCO lies in its role as a foundational tool for describing quantum systems: it provides a physically motivated framework for identifying all distinguishable states, predicting measurement outcomes, and ignoring irrelevant degrees of freedom, such as in atomic or particle physics experiments.^[2]^[3] For instance, in the hydrogen atom, the Hamiltonian \hat{H}, the square of the angular momentum \hat{L}^2, and its z-component \hat{L}_z form a CSCO, with eigenvalues labeled by quantum numbers n, l, and m_l, respectively, uniquely characterizing the energy eigenstates.^[1] A single non-degenerate observable, like the position operator in one dimension, can also constitute a CSCO by itself.^[3] In more complex systems, such as a free particle in three dimensions, the components of the momentum operator serve as a CSCO to resolve the state fully.^[3]

Background Concepts

Commuting Operators in Quantum Mechanics

In quantum mechanics, the commutator of two operators A and B, defined as [A, B] = AB - BA, quantifies whether their order of application affects the outcome when acting on a state vector. This algebraic structure emerged as a cornerstone of the theory, distinguishing quantum operations from their classical counterparts where multiplication is commutative. Observables represented by operators A and B are said to commute if [A, B] = 0, which implies that they can be measured simultaneously on the same quantum system with arbitrary precision, without mutual disturbance. This compatibility arises because the measurement outcomes remain consistent regardless of the sequence, reflecting an underlying joint probability distribution for their eigenvalues. The concept of commuting operators was introduced by Werner Heisenberg in his 1925 formulation of matrix mechanics, where non-commuting dynamical variables replaced classical trajectories to resolve atomic spectra inconsistencies. Paul Dirac further developed this in his 1927 transformation theory, emphasizing the commutator's role in deriving quantum equations from classical Poisson brackets via q \to q, p \to -i\hbar \frac{\partial}{\partial q}.^[4] A classic non-commuting pair is position x and momentum p, satisfying [x, p] = i\hbar, which underpins the Heisenberg uncertainty principle and prohibits precise simultaneous measurements. In contrast, in bound systems like the hydrogen atom, the total energy (Hamiltonian H) and parity operator P often commute, [H, P] = 0, allowing parity to label energy eigenstates without conflict. If two self-adjoint operators A and B commute, [A, B] = 0, they possess a common set of eigenstates, enabling a shared basis for state representation. This property, rooted in the spectral theorem for compatible observables, facilitates the labeling of quantum states by multiple commuting quantities.

Simultaneous Diagonalization and Common Eigenstates

In finite-dimensional Hilbert spaces, the spectral theorem asserts that every self-adjoint operator admits an orthonormal basis of eigenvectors with real eigenvalues, allowing it to be diagonalized via a unitary transformation.^[5] For a pair of commuting self-adjoint operators A and B on such a space, this extends to simultaneous diagonalization: there exists a unitary operator U such that both U^\dagger A U and U^\dagger B U are diagonal matrices, with the diagonal entries forming the joint spectrum of A and B.^[5] The columns of U constitute a common orthonormal basis of eigenvectors for A and B. This common basis arises because the commutator [A, B] = 0 implies that each eigenspace of A is invariant under B; thus, eigenvectors of A can be chosen to also diagonalize the restriction of B to that eigenspace.^[6] In the presence of degeneracy—where an eigenvalue of A has multiplicity greater than one—the commuting B acts within the degenerate subspace, enabling its eigenvalues to label and partially resolve the degeneracy by providing additional quantum numbers for the shared eigenstates.^[5] In infinite-dimensional separable Hilbert spaces, the result generalizes via von Neumann's spectral theorem, which represents bounded self-adjoint operators using commuting projection-valued measures when the operators commute. This joint spectral resolution allows a form of simultaneous "diagonalization" in the sense of multiplication operators on a product measure space, but caveats arise for unbounded operators or continuous spectra, where common eigenvectors may not span the space and generalized eigenfunctions (via rigged Hilbert spaces) are needed instead.^[6]

Formal Definition

Definition of a Complete Set

In quantum mechanics, a complete set of commuting observables (CSCO) is defined as a collection of Hermitian operators \{A_1, A_2, \dots, A_n\} that pairwise commute, i.e., [A_i, A_j] = 0 for all i, j, and whose simultaneous eigenstates form a complete orthonormal basis for the Hilbert space of the system.^[7] These common eigenstates, denoted in Dirac notation as |a_1, a_2, \dots, a_n\rangle, satisfy A_i |a_1, a_2, \dots, a_n\rangle = a_i |a_1, a_2, \dots, a_n\rangle for each i, where a_i are the corresponding eigenvalues.^[2] The set is complete when these eigenstates span the entire space without redundancy, ensuring that any state in the Hilbert space can be expanded in this basis. The joint eigenvalues (a_1, a_2, \dots, a_n) uniquely label each basis state, thereby resolving all degeneracies present in the individual operators.^[7] This uniqueness implies that no two distinct eigenstates share the same set of eigenvalues, providing a one-to-one correspondence between the labels and the states. The completeness of the basis is formalized by the resolution of the identity operator:

\sum_{a_1, a_2, \dots, a_n} |a_1, a_2, \dots, a_n\rangle \langle a_1, a_2, \dots, a_n| = \hat{I},

where the sum runs over all possible eigenvalue combinations, and \hat{I} is the identity operator on the Hilbert space.^[8] A CSCO is maximal among sets of commuting observables, meaning that adjoining any additional operator that commutes with all members of the set would either violate the completeness by introducing further degeneracy or render the new operator a function of the existing ones, failing to provide new independent information.^[2] This maximality ensures the set fully specifies the quantum state up to an overall phase.

Criteria for Completeness

A set of commuting observables \{A_1, A_2, \dots, A_n\} in a d-dimensional Hilbert space is complete if it satisfies specific theoretical and practical criteria that ensure its common eigenbasis fully spans the space without redundancy or unresolved degeneracies. These criteria verify maximality and uniqueness in labeling states, building on the formal requirement that the observables share a complete set of simultaneous eigenvectors. A related condition is that the number of distinct joint eigenvalues equals the Hilbert space dimension:

\text{Number of distinct } ( \lambda_1, \lambda_2, \dots, \lambda_n ) = d,

where \lambda_i are the eigenvalues of A_i. This equality confirms the set's sufficiency in distinguishing all states uniquely, with each common eigenspace being one-dimensional.^[1] Trace orthogonality provides another test: the joint eigenspaces must be one-dimensional, meaning the projection operators P_{\lambda_1, \dots, \lambda_n} onto each common eigenspace satisfy \operatorname{Tr}(P_{\lambda_1, \dots, \lambda_n}) = 1 and are mutually orthogonal, with \sum P_{\lambda_1, \dots, \lambda_n} = I. This orthogonality ensures the projectors form a resolution of the identity, verifying completeness by eliminating multi-dimensional degeneracies.^[9] Irreducibility, or maximality, requires that the set generates the full commutant algebra: any operator commuting with all A_i must be a function of the A_i themselves. This criterion, akin to the absence of non-trivial invariant subspaces (per Schur's lemma in representation theory), confirms the set cannot be extended by additional independent commuting observables. For systems with continuous spectra, such as position or momentum observables, completeness is assessed using rigged Hilbert spaces (Gel'fand triples), where improper eigenvectors replace square-integrable ones. Here, Dirac delta normalization \langle \phi | \psi \rangle = \delta(\lambda - \mu) ensures orthonormality. A single continuous observable, such as the position operator in one dimension, can form a CSCO, while incompatible pairs like position and momentum cannot be jointly complete due to non-commutation.^[9]

Key Theorems and Properties

The Compatibility Theorem

The compatibility theorem in quantum mechanics states that a set of pairwise commuting self-adjoint operators can be simultaneously diagonalized, meaning there exists a common orthonormal basis of simultaneous eigenvectors.^[1] This theorem is a consequence of the spectral theorem for self-adjoint operators and ensures that commuting observables share a common eigenbasis, allowing for their simultaneous measurement without disturbance. The theorem underpins the construction of complete sets of commuting observables (CSCOs). In particular, for a quantum system with self-adjoint Hamiltonian H, there exists a CSCO that includes H (or operators compatible with it), such that the common eigenstates uniquely label the degeneracy of the energy eigenspaces, fully specifying the state space.^[2] The concept was applied by Wolfgang Pauli in 1926 to analyze the hydrogen atom under external fields, where he identified commuting operators (such as the quantum Runge-Lenz vector) to resolve spectral degeneracies while preserving symmetries.^[10] It was rigorously integrated into the axiomatic foundations of quantum mechanics by John von Neumann in 1932, linking it to the spectral theory of self-adjoint operators. For a collection of bounded self-adjoint operators that pairwise commute, any such set can be extended to a maximal commuting set, meaning no additional bounded self-adjoint operator can commute with all members without being functionally dependent on them. This maximality criterion underpins the completeness of the set in the context of CSCOs. A central consequence is that if \{A_i\} forms a complete set of commuting observables, then any self-adjoint operator B that commutes with every A_i must be expressible as a Borel function of the A_i's, i.e., B = f(A_1, A_2, \dots, A_n) for some suitable f. The compatibility theorem is connected to more advanced results, such as the Stone–von Neumann theorem, which proves the uniqueness (up to unitary equivalence) of irreducible representations of the canonical commutation relations [q, p] = i\hbar in infinite-dimensional Hilbert space. This provides a foundational structure for systems like the harmonic oscillator, where commuting sets of ladder operators or number operators can be defined.^[11]

Proofs of the Theorem

The compatibility theorem states that a finite set of pairwise commuting self-adjoint operators on a finite-dimensional Hilbert space can be simultaneously diagonalized by a unitary transformation, meaning there exists an orthonormal basis consisting of common eigenvectors.^[12] The proof proceeds by induction on the dimension of the space and the number of operators. For a single self-adjoint operator A, the spectral theorem guarantees diagonalization in an orthonormal basis of eigenvectors. For a second commuting self-adjoint operator B, the spectral projections P_a onto the eigenspaces of A (corresponding to eigenvalue a) satisfy [A, B] = 0 implies P_a B = B P_a, so B preserves each eigenspace of A. Restricting B to each finite-dimensional eigenspace yields a self-adjoint operator that can be diagonalized within that subspace without affecting the eigenvalues of A, completing the step for two operators. Induction extends this to any finite set, as adding further commuting operators preserves the block-diagonal structure aligned with prior eigenspaces.^[12] To establish completeness of a commuting set \{A_i\}, consider the common eigenspaces labeled by joint eigenvalues (a_1, \dots, a_n). If any such space has dimension greater than 1, the set is partial, and extension is possible by constructing a self-adjoint operator B that commutes with all A_i but splits the degeneracy, for example, by diagonalizing a non-trivial Hermitian operator within that subspace.^[2] Completeness requires that all common eigenspaces are one-dimensional, ensuring unique labeling of basis states. This maximality follows from representation theory: the commutative algebra generated by the A_i acts on each common eigenspace, and since the representation is irreducible (by assumption of maximality), Schur's lemma implies that any operator commuting with all generators must be a scalar multiple of the identity on that space. For an abelian algebra, all irreducible representations are one-dimensional, confirming the eigenspaces cannot exceed dimension 1 without allowing further commuting extensions.^[13] The key commutation of projectors is formalized as follows: for commuting self-adjoint A and B, with spectral resolution A = \int \lambda \, dE(\lambda), the projector onto the eigenspace for eigenvalue a is P_a = E(\{a\}), and

\begin{equation} P_a B = B P_a, \end{equation}

which follows from the spectral theorem and [A, B] = 0 implying simultaneous refinement of spectral measures. This allows iterative resolution of degeneracies.^[12] In the infinite-dimensional case on a separable Hilbert space, the proof outlines a generalization using the joint spectral theorem for a commuting family of (possibly unbounded) self-adjoint operators, which decomposes the space as a direct integral over the joint spectrum \sigma(A_1, \dots, A_n) \subseteq \mathbb{R}^n:

\mathcal{H} \cong \int^\oplus_{\sigma(A_1, \dots, A_n)} \mathcal{H}(\lambda) \, d\mu(\lambda),

where each fiber \mathcal{H}(\lambda) carries the multiplicity of the joint "eigenvalue" \lambda, and the operators act by multiplication by coordinate functions on the spectrum.^[14] A set is complete if the multiplicity is 1 almost everywhere with respect to the spectral measure \mu, ensuring no further commuting self-adjoint extensions exist to resolve degeneracies. The Gelfand-Naimark theorem underpins this by guaranteeing that the C*-algebra generated by the bounded functions of the operators admits a faithful *-representation on the Hilbert space, facilitating the direct integral construction in the separable case.^[15] These proofs assume all operators are self-adjoint to invoke the spectral theorem, ensuring real spectra and unitary diagonalizability, and rely on the separability of the Hilbert space to guarantee the existence of a complete (generalized) eigenbasis, even in the presence of continuous spectrum where eigenvectors are replaced by spectral projections.^[14]

Implications for Measurement

The presence of a complete set of commuting observables (CSCO) enables the simultaneous measurement of all operators in the set, as their mutual commutativity ensures compatible measurement outcomes without inherent trade-offs in precision. This compatibility allows for the determination of a precise joint probability distribution for the eigenvalues of these observables, free from the constraints imposed by uncertainty relations on non-commuting pairs. For a quantum state |\psi\rangle, the probability P(a_1, \dots, a_n) of obtaining eigenvalues a_1, \dots, a_n upon joint measurement of the CSCO operators \hat{A}_1, \dots, \hat{A}_n is given by

P(a_1, \dots, a_n) = |\langle a_1 \dots a_n | \psi \rangle|^2,

where |a_1 \dots a_n \rangle denote the common eigenstates that form an orthonormal basis for the Hilbert space. The common eigenstates of a CSCO uniquely label quantum states through their eigenvalues, which serve as a complete set of quantum numbers specifying the state without degeneracy in the labeling. This unique assignment is fundamental for theoretical frameworks like time-independent perturbation theory, where unperturbed eigenstates are identified by these quantum numbers to systematically compute corrections to energies and wavefunctions under small perturbations. Since membership in a CSCO requires all operators to commute pairwise, non-commuting observables are excluded from the same set, aligning directly with the uncertainty principle: the standard deviation in measurements of two observables satisfies \Delta A \Delta B \geq \frac{1}{2} |\langle [\hat{A}, \hat{B}] \rangle|, which is non-zero for non-commuting pairs and thus precludes their joint precise determination. In atomic physics experiments, CSCOs underpin the derivation of selection rules for radiative transitions, as the matrix elements of interaction operators between states labeled by CSCO eigenvalues dictate allowed changes in quantum numbers. For example, in hydrogen-like atoms, the CSCO {Hamiltonian, \hat{L}^2, \hat{L}_z} labels states as |n, l, m\rangle, leading to electric dipole selection rules such as \Delta l = \pm 1 and \Delta m = 0, \pm 1, which govern observable spectral line intensities and positions.

Examples and Applications

Hydrogen Atom without Spin

The non-relativistic Hamiltonian for the hydrogen atom, excluding electron spin, is given by

H = \frac{\mathbf{p}^2}{2m} - \frac{e^2}{r},

where \mathbf{p} is the electron momentum operator, m is the electron mass, e is the elementary charge, and r = |\mathbf{r}| is the distance from the proton.^[16] This operator governs the time-independent Schrödinger equation H \psi = E \psi for bound states.^[17] The angular momentum operators satisfy [L_x, L_y] = i \hbar L_z (and cyclic permutations), with L^2 = L_x^2 + L_y^2 + L_z^2 and L_z = -i \hbar \partial / \partial \phi in spherical coordinates.^[18] Crucially, the Hamiltonian commutes with both L^2 and L_z:

[H, L^2] = 0, \quad [H, L_z] = 0,

and inherently [L^2, L_z] = 0.^[17] These relations follow from the central (spherically symmetric) form of the Coulomb potential, preserving rotational invariance. Consequently, the set \{H, L^2, L_z\} constitutes a complete set of commuting observables (CSCO), as defined by a maximal collection of mutually commuting Hermitian operators whose common eigenbasis fully spans the Hilbert space. The common eigenstates of this CSCO are denoted |n, l, m\rangle, where n = 1, 2, \dots is the principal quantum number determining the energy E_n = - (m e^4)/(2 \hbar^2 n^2), l = 0, 1, \dots, n-1 labels eigenvalues of L^2 via \hbar^2 l(l+1), and m = -l, \dots, l labels eigenvalues of L_z via \hbar m.^[17] This set resolves the n^2-fold degeneracy of each energy level, as the number of states sums over l from 0 to n-1 and m from -l to l, yielding exactly n^2 orthogonal states per E_n.^[16] The wave functions factorize as \psi_{n l m}(r, \theta, \phi) = R_{n l}(r) Y_{l m}(\theta, \phi), with radial parts R_{n l} solving an effective one-dimensional equation and angular parts as spherical harmonics.^[17] The CSCO enables separation of variables in the Schrödinger equation under spherical coordinates, transforming the partial differential equation into independent ordinary differential equations for the radial function (involving an effective potential with centrifugal term \hbar^2 l(l+1)/(2 m r^2)) and the angular functions (yielding the spherical harmonics as eigenfunctions of L^2 and L_z).^[16] This separability, rooted in the commuting observables, yields exact analytic solutions for all bound states, confirming the discrete spectrum and degeneracy structure.

Free Particle in One Dimension

The free particle in one dimension provides a foundational example of a complete set of commuting observables (CSCO) in quantum mechanics, where the absence of a potential leads to translational invariance and a continuous spectrum of energies. The Hamiltonian for this system is given by H = \frac{p^2}{2m}, where p is the momentum operator and m is the particle mass, with no external potential V(x) = 0.^[19] Since the momentum is conserved under translation, the operators H and p commute, [H, p] = 0, forming a CSCO whose common eigenstates uniquely label the quantum states.^[19] These eigenstates are the momentum eigenstates |p\rangle, satisfying H |p\rangle = E_p |p\rangle and p |p\rangle = p |p\rangle, with energy eigenvalue E_p = \frac{p^2}{2m}. In contrast to bound systems like the hydrogen atom, which exhibit discrete energy levels, the free particle's spectrum is continuous for E \geq 0, reflecting unbounded motion. The momentum eigenstates |p\rangle form a complete basis for the Hilbert space due to the continuous nature of the spectrum, with orthonormality enforced by the [Dirac delta function](/page/Dirac delta function): \langle p | p' \rangle = \delta(p - p').^[19] This normalization ensures completeness via the resolution of the identity \int_{-\infty}^{\infty} dp \, |p\rangle \langle p| = \hat{I}, allowing any state |\psi\rangle to be expanded as |\psi\rangle = \int_{-\infty}^{\infty} dp \, \psi(p) |p\rangle, where \psi(p) = \langle p | \psi \rangle. In the position representation, the corresponding wavefunctions are plane waves: \psi_p(x) = \langle x | p \rangle = \frac{1}{\sqrt{2\pi \hbar}} e^{i p x / \hbar}.^[19] For each energy E > 0, there is a twofold degeneracy corresponding to p = \pm \sqrt{2 m E}, but specifying p uniquely resolves this, making \{p\} itself a CSCO sufficient for labeling states without further operators. The continuous degeneracy and spectrum pose challenges in practical treatments, as the Dirac delta normalization is idealized for infinite space. To approximate discrete states, one often imposes periodic boundary conditions in a large box of length L, discretizing the momentum as p_n = \frac{2\pi \hbar n}{L} for integer n, with orthonormal plane waves \psi_n(x) = \frac{1}{\sqrt{L}} e^{i p_n x / \hbar} and \langle n | n' \rangle = \delta_{n n'}.^[19] In the limit L \to \infty, this recovers the continuous case, facilitating calculations like scattering or wave packet evolution while preserving the CSCO structure. This approach highlights how CSCOs adapt to continuous systems, enabling precise predictions of observables like position and momentum uncertainties through Fourier transforms between representations.

Addition of Angular Momenta

In quantum mechanics, the addition of angular momenta is essential for describing composite systems, such as two particles with individual angular momentum operators \mathbf{J_1} and \mathbf{J_2}, where the total angular momentum operator is defined as \mathbf{J} = \mathbf{J_1} + \mathbf{J_2}. This construction preserves the algebraic structure of angular momentum, with \mathbf{J} satisfying the same commutation relations as a single angular momentum operator. For such a system, a complete set of commuting observables (CSCO) consists of \{ \mathbf{J}^2, J_z, \mathbf{J_1}^2, \mathbf{J_2}^2 \}, whose common eigenstates are denoted |j, m; j_1, j_2 \rangle. Here, j_1 and j_2 are the fixed eigenvalues of \sqrt{\mathbf{J_1}^2} and \sqrt{\mathbf{J_2}^2} (in units of \hbar), while j ranges from |j_1 - j_2| to j_1 + j_2 in integer steps, and m takes values from -j to +j. These labels uniquely specify the states within the relevant Hilbert subspace. The operators in this CSCO commute pairwise: [ \mathbf{J}^2, J_z ] = 0, [ \mathbf{J_1}^2, \mathbf{J}^2 ] = 0, [ \mathbf{J_2}^2, \mathbf{J}^2 ] = 0, [ \mathbf{J_1}^2, J_z ] = 0, [ \mathbf{J_2}^2, J_z ] = 0, and [ \mathbf{J_1}^2, \mathbf{J_2}^2 ] = 0, allowing simultaneous measurement and diagonalization. The completeness arises because the degeneracy structure matches the subspace dimension: the total number of states is (2j_1 + 1)(2j_2 + 1), and summing the multiplicities (2j + 1) over allowed j exactly reproduces this value, ensuring no further commuting observables are needed to resolve degeneracies.^[20] The coupled basis |j, m; j_1, j_2 \rangle relates to the uncoupled product basis |j_1, m_1; j_2, m_2 \rangle (eigenstates of \mathbf{J_1}^2, J_{1z}, \mathbf{J_2}^2, J_{2z}) via a unitary transformation involving Clebsch-Gordan coefficients:

|j, m; j_1, j_2 \rangle = \sum_{m_1, m_2} \langle j_1 m_1 j_2 m_2 | j m \rangle \, |j_1, m_1; j_2, m_2 \rangle,

where m = m_1 + m_2 and the coefficients \langle j_1 m_1 j_2 m_2 | j m \rangle are real numbers determined by the theory, often tabulated for practical use. These coefficients encode the recoupling and ensure orthogonality in the coupled representation. This approach is used in constructing the total angular momentum for multi-electron atoms by coupling individual electron angular momenta, but due to electron indistinguishability, the states are classified using total quantum numbers L, S, J in the LS coupling scheme, with individual l values specified by the electron configuration (e.g., 2p^2). The CSCO typically includes {H, \mathbf{L}^2, L_z, \mathbf{S}^2, S_z} (ignoring spin-orbit) for energy level classification, extending the hydrogenic framework while accounting for electron interactions and symmetry.^[21]

Advanced Extensions

Maximal Commuting Sets

A maximal commuting set of observables in quantum mechanics is defined as a set of self-adjoint operators \{A_i\} that pairwise commute and cannot be extended by including another self-adjoint operator B that commutes with all A_i unless B is a function of the A_i, thereby ensuring irreducibility in the sense that the set fully specifies the common eigenspaces without redundancy.^[22] In the algebraic formulation of quantum mechanics, such maximal sets generate maximal abelian von Neumann subalgebras (MASAs), which correspond to Cartan subalgebras in the associated Lie algebra structure, particularly when considering the Weyl algebra defined by canonical commutation relations; this equivalence arises because the Cartan subalgebra consists of mutually commuting, diagonalizable elements that maximize the number of simultaneously measurable observables.^[23]^[11] A fundamental result, extending von Neumann's double commutant theorem, states that every set of commuting observables can be enlarged to a maximal commuting set by iteratively adding operators that preserve commutativity, leveraging Zorn's lemma in the lattice of abelian subalgebras; this guarantees the existence of such maximal extensions in the bounded operators on a Hilbert space. For a maximal commuting set \{A_i\}, the commutant, defined as \{B \mid [B, A_i] = 0 \ \forall i\}, coincides with the von Neumann algebra generated by polynomials in the A_i:

\{B \mid [B, A_i] = 0 \ \forall i\} = \text{vN}\{p(A_1, \dots, A_n) \mid p \text{ polynomial}\},

where vN denotes the weak closure, ensuring that all operators commuting with the set are functionally dependent on it.^[22] In infinite-dimensional Hilbert spaces, maximal commuting sets exhibit non-uniqueness, as distinct MASAs may not be unitarily conjugate, leading to pathological cases where multiple non-equivalent choices exist for the same system, unlike the finite-dimensional scenario where the structure is more rigid; this underemphasis in standard treatments highlights the subtleties of unbounded operators and spectral theory in continuous spectra.^[24]

Relation to Symmetry Groups

In quantum mechanics, symmetry groups are represented by unitary operators acting on the Hilbert space, preserving the inner product and probabilities as per Wigner's theorem. These unitary representations of Lie groups generate sets of commuting observables through their Lie algebra generators and invariant elements known as Casimir operators. Casimir operators, constructed from polynomials in the generators, commute with all group elements and thus with every generator G_a, satisfying [C, G_a] = 0 for all a. This property allows them to form part of a complete set of commuting observables (CSCO), labeling the irreducible representations (irreps) of the group and enabling the classification of quantum states invariant under the symmetry.^[13]^[25] A canonical example arises in the SO(3) rotation group, which governs angular momentum in quantum systems with spherical symmetry. The generators are the angular momentum components \mathbf{L} = (L_x, L_y, L_z), satisfying the commutation relations [L_i, L_j] = i \hbar \epsilon_{ijk} L_k. The Casimir operator L^2 = L_x^2 + L_y^2 + L_z^2 commutes with all generators, including L_z, forming the CSCO \{L^2, L_z\}. Their common eigenstates |l, m\rangle transform irreducibly under SO(3), with eigenvalues L^2 |l, m\rangle = \hbar^2 l(l+1) |l, m\rangle and L_z |l, m\rangle = \hbar m |l, m\rangle, where l = 0, 1/2, 1, \dots and m = -l, \dots, l. This set uniquely labels states in irreps of dimension $2l + 1.^[26]^[13] Wigner's theorem guarantees that symmetry transformations correspond to unitary (or anti-unitary) operators on the Hilbert space, ensuring that CSCOs consisting of Casimirs and Cartan subalgebra generators label the irreps of the symmetry group. The common eigenstates of such a CSCO transform irreducibly under the group action, spanning the representation space without invariant subspaces. By Schur's lemma, any operator commuting with all group elements in an irrep must be a multiple of the identity, reinforcing the maximality of the CSCO.^[25] In quantum field theory, this framework extends to the Poincaré group, the symmetry group of special relativity. The Casimir operators P^\mu P_\mu = m^2 (mass squared) and W^\mu W_\mu = -m^2 s(s+1) (Pauli-Lubanski squared, related to spin s), commute with all Poincaré generators and form a CSCO for one-particle states. These invariants classify elementary particles by mass and spin, with eigenstates transforming irreducibly under the little group (e.g., SO(3) for massive particles). This addresses the labeling of relativistic quantum states beyond non-relativistic symmetries.^[27]

References

[1]
[PDF] Observables
In this case, the observable A constitutes by itself a complete set of commuting observables. (CSCO), i.e. the eigenvalues of ˆA are sufficient to identify the ...
[2]
None
### Summary of Complete Set of Commuting Observables (CSCO) from the Document
[3]
[PDF] Lecture 14 CSCOs, Energy-Time Uncertainty, and Quasi-Bound States
Why do we care about CSCOs? 1) It is a physically motivated assumption that for any physical quantum system there is a complete set of commuting observables.
[4]
[PDF] Spectral theory in Hilbert spaces (ETH Zürich, FS 09) E. Kowalski
We can now prove Theorem 3.1 for self-adjoint operators. Theorem 3.17 (Spectral theorem for self-adjoint operators). Let H be a separable. Hilbert space and ...
[5]
https://people.math.ethz.ch/~kowalski/spectral-theory.pdf
[6]
[PDF] UNCERTAINTY PRINCIPLE AND COMPATIBLE OBSERVABLES
Oct 21, 2013 · In this case the operator S provides a “complete set of commuting observables” or a CSCO, in short. The set here has just one observable, the ...
[7]
[PDF] Quantum Physics
Sep 15, 2012 · 5.7.1 Complete set of commuting observables . ... by the completeness relation I = Pj′′,m′′,α′′ |j′′,m′′,α′′ihj′′,m ...
[8]
[PDF] Quantum Mechanics Charles B. Thorn1 - UF Physics
We then need more observables to make this distinction. A Complete Set of Commuting observables is a set of mutually commuting observables for which there ...
[9]
[PDF] The hydrogen atom in electric and magnetic fields : Pauli's 1926 article
that the complete set of commuting observables diagonalized here are H0, Iz, and Kz while in. Schrödinger's approach one takes H0, L 2, and Lz. Note also ...
[10]
[PDF] A Selective History of the Stone-von Neumann Theorem - UMD MATH
Abstract. The names of Stone and von Neumann are intertwined in what is now known as the Stone-von Neumann Theorem. We discuss the origins of.
[11]
[PDF] Quantum Theory I, Recitation 1 Notes - MIT OpenCourseWare
Thus, A and B are simultaneously diagonalizable. This proves the backward direction: if A and B commute, then they are simultaneously diagonalizable.
[12]
[PDF] Quantum Theory, Groups and Representations: An Introduction ...
“Quantization” is then the passage to a unitary representation (unique by the Stone-von Neumann theorem) of a subalgebra of this Lie algebra. • The role of the ...
[13]
[PDF] Spectral Theorem for Commuting Normal Operators - UBC Math
Nov 19, 2018 · Both Re A and ImA are self–adjoint. The operator A is normal if and only if Re A and ImA commute. In these notes we prove. Theorem 3 (Spectral ...
[14]
Gelfand-Naimark theorem | What's new - Terry Tao
Feb 10, 2010 · ... infinite-dimensional setting of the large {n} limit is non-pathological. (This is ultimately due to the stable nature of eigenvalues in the ...
[15]
[PDF] 3. Quantisation as an eigenvalue problem; by E. Schrödinger
∗Original title: Quantisierung als Eigenwertproblem. Published in: Annalen der Physik 79. (1926): 361-376. Translated by Oliver F. Piattella. E-mail: oliver ...
[16]
[PDF] QUANTUM MECHANICS-Non-Relativistic Theory
The topics dealt with include the basic concepts, Schrodinger's equation, angular momentum, motion in centrally symmetric fields perturbation theory, the quasi- ...
[17]
https://power1.pc.uec.ac.jp/~toru/notes/LandauLifshitz-QuantumMechanics.pdf
[18]
[PDF] Quantum Mechanics - Northern Illinois University
Apr 2, 2025 · ... free-particle state in one dimension ... Complete Set of Commuting Observables, called a. CSCO from here on. For any Hilbert space of states ...<|separator|>
[19]
[PDF] ANGULAR MOMENTUM - MIT OpenCourseWare
Dec 16, 2013 · We can now consider the issue of complete sets of commuting observables. The list of operators that we have is. H, x1,x2,x3, p1,p2,p3, L1,L2 ...
[20]
https://ocw.mit.edu/courses/8-05-quantum-physics-ii-fall-2013/e765f050b2fd50a9d87d2ae801e9c52f_MIT8_05F13_Chap_09.pdf
[21]
Cartan Subalgebra Approach to Efficient Measurements of Quantum Observables
### Definition of Cartan Subalgebra in Context of Quantum Observables and Commuting Sets
[22]
[PDF] New Journal of Physics - University of Waterloo
Jun 22, 2009 · formalism of infinite-dimensional quantum theory. The operators satisfy the canonical commutation relations. [Q, P] = i. Since Q and P do not ...
[23]
[PDF] Symmetries in Quantum Mechanics: From Angular Momentum to ...
) in which no super-selection box structure remains. 2.4.4 Complete symmetry group, complete sets of commuting observables, complete sets of states. The ...
[24]
[PDF] Chapter 9 Angular Momentum Quantum Mechanical Angular ...
We have indicated H, L2, and Lz form a complete set of commuting observables for such a system. You may be familiar with the principal quantum number n, the ...
[25]
[PDF] Poincaré group - CFTP
Here lies the origin of spin: from the point of view of the Poincaré group, the mass m and spin j are the only Poincaré-invariant quantum numbers that we can.