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Quantum harmonic oscillator

The quantum harmonic oscillator is a foundational model in quantum mechanics that describes the behavior of a particle confined by a restoring force proportional to its displacement from an equilibrium position, analogous to a classical mass-spring system but with quantized energy states.^[1] This system arises from solving the time-independent Schrödinger equation for a potential energy given by V(x) = \frac{1}{2} m \omega^2 x^2, where m is the particle's mass, \omega is the angular frequency, and x is the displacement.^[2] The energy eigenvalues are discrete and evenly spaced, expressed as E_n = \hbar \omega \left(n + \frac{1}{2}\right) for n = 0, 1, 2, \dots, with \hbar denoting the reduced Planck's constant, resulting in a non-zero ground-state energy known as zero-point energy.^[3] The corresponding wavefunctions are stationary states that exhibit oscillatory probability densities, involving Hermite polynomials multiplied by a Gaussian factor, and they form a complete orthonormal basis for one-dimensional quantum systems.^[4] This exact solvability makes the quantum harmonic oscillator a cornerstone for introducing key quantum concepts such as ladder operators (creation and annihilation operators), which simplify the algebraic treatment of the spectrum and coherent states.^[5] Beyond its theoretical elegance, the model has broad applications in physics, serving as an approximation for small-amplitude vibrations in molecules, phonons in solids, electromagnetic modes in cavities, and even superconducting circuits in quantum computing.^[6] It underpins the quantum theory of radiation, the description of blackbody radiation via Planck's law, and the quantization of fields in quantum field theory, where fields are expanded in terms of harmonic oscillator modes.^[7]

Classical Background

Classical Harmonic Oscillator

The classical harmonic oscillator models a mass-spring system where a particle of mass m attached to a spring experiences a restoring force proportional to its displacement x from equilibrium, as described by Hooke's law: F = -kx, with k > 0 the spring constant.^[8] This linear force law leads to periodic motion and serves as a foundational approximation for many oscillatory phenomena in physics.^[8] The concept originates from Robert Hooke's 1678 publication Lectures de potentia restitutiva, or of spring, which explained the elastic properties of springing bodies.^[9] Applying Newton's second law yields the second-order differential equation governing the motion:

m \frac{d^2 x}{dt^2} + kx = 0,

or equivalently \ddot{x} + \omega^2 x = 0, where \omega = \sqrt{k/m} is the natural angular frequency.^[8] The general solution is a sinusoidal oscillation:

x(t) = A \cos(\omega t + \phi),

with amplitude A and phase \phi determined by initial conditions.^[8] This solution describes simple harmonic motion, where the period T = 2\pi / \omega depends only on the intrinsic properties k and m, independent of amplitude for small displacements.^[8] The total mechanical energy E is conserved due to the absence of dissipation, expressed as E = \frac{1}{2} k A^2 = \frac{1}{2} m \omega^2 A^2.^[8] It partitions between kinetic energy \frac{1}{2} m \dot{x}^2 and potential energy \frac{1}{2} k x^2, oscillating between maximum kinetic (at x = 0) and maximum potential (at x = \pm A) values.^[8] In phase space, plotting position x against momentum p = m \dot{x}, the trajectories form closed elliptical orbits enclosing constant energy, with the area of each ellipse proportional to E.^[10] Hooke's model provides the basis for approximating other systems, such as the simple pendulum, where for small angular displacements \theta, the restoring torque yields an effective force F \approx -(mg/L) s (with arc length s = L \theta), mimicking F = -kx with k = mg/L and frequency \omega = \sqrt{g/L}.^[11]

Transition to Quantum Mechanics

The transition from the classical to the quantum description of the harmonic oscillator began within the framework of old quantum theory, where Niels Bohr and Arnold Sommerfeld introduced semiclassical quantization rules to reconcile atomic spectra with classical mechanics. In 1913, Bohr proposed that the angular momentum in atomic orbits is quantized in units of Planck's constant h, but for periodic systems like the harmonic oscillator, Sommerfeld generalized this in 1916 to the action integral over a closed orbit: \oint p \, dq = n h, where n is a non-negative integer, p is momentum, and q is the coordinate.^[12] For the harmonic oscillator, this rule yields the energy levels E_n = n \hbar \omega, where \hbar = h / 2\pi and \omega is the classical angular frequency; this approximates the excited states but misses the zero-point energy, unlike the exact quantum result E_n = \left(n + \frac{1}{2}\right) \hbar \omega. The full quantum mechanical treatment emerged in 1926 with Erwin Schrödinger's wave mechanics, where the classical Hamiltonian H = \frac{p^2}{2m} + \frac{1}{2} m \omega^2 x^2 is promoted to an operator form \hat{H} = \frac{\hat{p}^2}{2m} + \frac{1}{2} m \omega^2 \hat{x}^2, with position \hat{x} and momentum \hat{p} satisfying the canonical commutation relation [\hat{x}, \hat{p}] = i \hbar.^[13] The time-independent Schrödinger equation \hat{H} \psi = E \psi then governs the energy eigenfunctions \psi, replacing the deterministic classical trajectories with probabilistic wave functions whose squared modulus gives the position probability density. This operator formulation resolves inconsistencies in old quantum theory by incorporating non-commuting observables inherently.^[13] Bohr's correspondence principle, articulated in 1923, ensures continuity between quantum and classical regimes: as \hbar \to 0 or for large quantum numbers n, the quantum harmonic oscillator's expectation values and transition probabilities approach classical periodic motion and radiation. For high n, the quantum probability distribution |\psi_n(x)|^2 localizes around the classical turning points, mimicking the classical energy equipartition between kinetic and potential forms. The Heisenberg uncertainty principle, formulated in 1927, further underscores the quantum departure from classical determinism, stating that the product of position and momentum spreads satisfies \Delta x \Delta p \geq \frac{\hbar}{2}. In the quantum harmonic oscillator, the ground state achieves this minimum uncertainty, with \Delta x = \sqrt{\frac{\hbar}{2 m \omega}} and \Delta p = \sqrt{\frac{m \hbar \omega}{2}}, implying unavoidable fluctuations even at zero-point energy E_0 = \frac{1}{2} \hbar \omega. Higher excited states exhibit larger spreads, but the principle prohibits simultaneous classical precision in both variables, leading to oscillatory probability distributions rather than definite trajectories.

One-Dimensional Quantum Harmonic Oscillator

Hamiltonian and Schrödinger Equation

The quantum harmonic oscillator serves as a foundational model in quantum mechanics, representing systems confined by a parabolic potential, such as vibrational modes in molecules or trapped particles.^[14] This formulation draws a direct analogy to the classical harmonic oscillator, where the total energy is expressed through the Hamiltonian combining kinetic and potential terms./University_Physics_III_-Optics_and_Modern_Physics(OpenStax)/07%3A_Quantum_Mechanics/7.06%3A_The_Quantum_Harmonic_Oscillator) In the quantum description, the Hamiltonian operator for the one-dimensional case is

\hat{H} = \frac{\hat{p}^2}{2m} + \frac{1}{2} m \omega^2 \hat{x}^2,

where m is the mass of the particle, \omega is the angular frequency of oscillation, \hat{x} is the position operator (multiplication by the coordinate x), and \hat{p} = -i \hbar \frac{d}{dx} is the momentum operator, with \hbar denoting the reduced Planck's constant.^[14] The potential energy term V(x) = \frac{1}{2} m \omega^2 x^2 describes the parabolic confinement that binds the particle, ensuring discrete energy levels for the system./University_Physics_III_-Optics_and_Modern_Physics(OpenStax)/07%3A_Quantum_Mechanics/7.06%3A_The_Quantum_Harmonic_Oscillator) To find the stationary states, one solves the time-independent Schrödinger equation

\hat{H} \psi(x) = E \psi(x),

where \psi(x) is the wave function and E is the energy eigenvalue.^[14] For bound states, the wave functions must satisfy the boundary condition \psi(x) \to 0 as |x| \to \infty, ensuring normalizability and physical relevance in the infinite parabolic well. The parameters m, \omega, and \hbar define the natural scales of length, time, and energy for the system, with characteristic length \sqrt{\hbar / m \omega} and energy \hbar \omega./University_Physics_III_-Optics_and_Modern_Physics(OpenStax)/07%3A_Quantum_Mechanics/7.06%3A_The_Quantum_Harmonic_Oscillator)

Energy Eigenvalues and Eigenfunctions

The time-independent Schrödinger equation for the one-dimensional quantum harmonic oscillator, as formulated in the previous section, admits exact analytical solutions through separation of variables and reduction to a known differential equation form.^[13] By introducing the dimensionless variable \xi = \sqrt{m \omega / \hbar} \, x, the equation transforms into the standard form \frac{d^2 \psi}{d \xi^2} + (2\nu + 1 - \xi^2) \psi = 0, where \nu = 2E / \hbar \omega - 1.^[13] For the solutions to remain normalizable (i.e., square-integrable), \nu must be a non-negative integer n = 0, 1, 2, \dots, yielding the energy eigenvalues E_n = \hbar \omega \left( n + \frac{1}{2} \right).^[13] This discrete spectrum reflects the quantization of energy, with the ground state (n=0) possessing a non-zero zero-point energy of E_0 = \frac{1}{2} \hbar \omega, even at absolute zero temperature.^[13] The corresponding energy eigenfunctions are expressed in terms of Hermite polynomials H_n(\xi), defined recursively or via the Rodrigues formula H_n(\xi) = (-1)^n e^{\xi^2} \frac{d^n}{d \xi^n} e^{-\xi^2}, as

\psi_n(x) = N_n \, H_n(\xi) \, e^{-\xi^2 / 2},

where N_n is the normalization constant ensuring \int_{-\infty}^{\infty} |\psi_n(x)|^2 \, dx = 1.^[13] The explicit form of the normalization is N_n = \left( \frac{m \omega}{\pi \hbar} \right)^{1/4} \frac{1}{\sqrt{2^n n!}}, derived from the orthogonality of Hermite polynomials \int_{-\infty}^{\infty} H_m(\xi) H_n(\xi) e^{-\xi^2} \, d\xi = \sqrt{\pi} \, 2^n n! \, \delta_{mn}./03%3A_Mostly_1-D_Quantum_Mechanics/3.04%3A_The_Simple_Harmonic_Oscillator) These wavefunctions form a complete orthonormal basis for the Hilbert space of the system.^[13] The eigenfunctions exhibit definite parity: \psi_n(-x) = (-1)^n \psi_n(x), with even n yielding even functions and odd n yielding odd functions, a consequence of the parity invariance of the harmonic potential./03%3A_Mostly_1-D_Quantum_Mechanics/3.04%3A_The_Simple_Harmonic_Oscillator) For the ground state, H_0(\xi) = [1](/page/1), so \psi_0(x) = \left( \frac{m \omega}{\pi [\hbar](/page/H-bar)} \right)^{[1](/page/1)/4} e^{-\xi^2 / 2}, a Gaussian centered at x = 0 with no nodes. Higher states introduce n nodes, with the oscillatory behavior of H_n(\xi) confined within the Gaussian envelope./03%3A_Mostly_1-D_Quantum_Mechanics/3.04%3A_The_Simple_Harmonic_Oscillator) The probability density |\psi_n(x)|^2 for large n approximates the classical probability distribution for a harmonic oscillator of energy E_n, peaking near the classical turning points x = \pm \sqrt{2 E_n / m \omega^2}, where the classical particle spends the most time. This correspondence principle holds as n \to \infty, with quantum tunneling allowing small probabilities beyond the turning points, unlike the classical case.

Ladder Operator Method

The ladder operator method offers an elegant algebraic framework for determining the energy eigenvalues and eigenstates of the quantum harmonic oscillator, bypassing the need to solve the differential Schrödinger equation directly. Introduced by Paul Dirac in his seminal work on quantum mechanics, this approach leverages non-commuting operators to reveal the discrete spectrum and structure of the states.^[15] Central to this formalism are the lowering operator a and the raising operator a^\dagger, defined in terms of the position operator x and momentum operator p as

a = \sqrt{\frac{m\omega}{2\hbar}} \left( x + \frac{i p}{m \omega} \right), \quad a^\dagger = \sqrt{\frac{m\omega}{2\hbar}} \left( x - \frac{i p}{m \omega} \right),

where m is the particle mass, \omega is the angular frequency, and \hbar is the reduced Planck's constant. These operators satisfy the canonical commutation relation [a, a^\dagger] = 1, which follows from the fundamental [x, p] = i\hbar.^[16] The number operator is given by N = a^\dagger a, which counts the excitations in the system and has non-negative integer eigenvalues. The Hamiltonian of the quantum harmonic oscillator can then be expressed compactly as \hat{H} = \hbar \omega \left( N + \frac{1}{2} \right), where the zero-point energy \frac{1}{2} \hbar \omega arises naturally from this representation.^[16] The energy eigenstates |n\rangle (with n = 0, 1, 2, \dots) satisfy \hat{H} |n\rangle = E_n |n\rangle, where E_n = \hbar \omega \left( n + \frac{1}{2} \right), and the ladder operators act to connect these states: a |n\rangle = \sqrt{n} |n-1\rangle and a^\dagger |n\rangle = \sqrt{n+1} |n+1\rangle. This generates an infinite ladder of states starting from the ground state |0\rangle, defined by the annihilation condition a |0\rangle = 0, ensuring no states exist below it. Higher states are constructed recursively as |n\rangle = \frac{(a^\dagger)^n}{\sqrt{n!}} |0\rangle.^[16] In the position representation, the ground state wavefunction is

\psi_0(x) = \left( \frac{m \omega}{\pi \hbar} \right)^{1/4} \exp\left( -\frac{m \omega x^2}{2 \hbar} \right),

a Gaussian form that normalizes to unity and minimizes the uncertainty product under the harmonic potential. This algebraic method yields the same energy eigenvalues as the direct solution of the Schrödinger equation while highlighting the oscillator's connections to broader quantum systems like angular momentum.^[16]

Natural Length and Energy Scales

The quantum harmonic oscillator introduces natural length and energy scales that facilitate the analysis of the system and provide insight into the quantum-to-classical transition. The characteristic length scale is defined as x_0 = \sqrt{\frac{[\hbar](/page/H-bar)}{m [\omega](/page/Angular_frequency)}}, where m is the mass of the particle, \omega is the angular frequency, and \hbar is the reduced Planck's constant. This scale serves as the quantum analog of the classical oscillation amplitude and arises in the exact solution of the time-independent Schrödinger equation for the potential V(x) = \frac{1}{2} m \omega^2 x^2.^[17] The characteristic energy scale is \hbar \omega, which sets the uniform spacing between the discrete energy levels of the oscillator. These levels are given by E_n = \hbar \omega \left( n + \frac{1}{2} \right) for non-negative integers n, reflecting the quantized vibrational modes inherent to the quantum description. This energy quantization was first established through the eigenvalue approach to the Schrödinger equation.^[17]^[18] By introducing the dimensionless position variable \xi = x / x_0, the Schrödinger equation can be recast in a parameter-free form, simplifying both analytical and numerical treatments. This scaling reveals the intrinsic structure of the problem, independent of specific physical units. In the ground state (n = 0), the wave function spreads over a width proportional to x_0, embodying the inherent quantum delocalization, while the product of position and momentum uncertainties achieves the minimum value \Delta x \Delta p = \frac{\hbar}{2} mandated by the uncertainty principle.^[18] In the semiclassical regime of large n, the quantum oscillator's behavior aligns with classical expectations. The classical turning points, or effective amplitude, for energy E_n scale as \sqrt{\frac{2 E_n}{m \omega^2}} \approx \sqrt{2n} \, x_0, demonstrating how the characteristic length x_0 governs the relative size of quantum fluctuations against the growing classical excursion.^[3]

Special States and Representations

Coherent States

Coherent states of the quantum harmonic oscillator are minimum-uncertainty Gaussian wave packets that exhibit classical-like oscillatory motion while preserving the minimum Heisenberg uncertainty product in position and momentum. These states were originally constructed by Erwin Schrödinger in 1926 to demonstrate a smooth transition from quantum to classical descriptions of the oscillator, where the wave packet centers follow the classical trajectory without spreading. The coherent state |\alpha\rangle, labeled by a complex parameter \alpha, is defined in the number basis as the normalized superposition

|\alpha\rangle = e^{-|\alpha|^2/2} \sum_{n=0}^{\infty} \frac{\alpha^n}{\sqrt{n!}} |n\rangle,

where |n\rangle are the energy eigenstates. Equivalently, it can be expressed as the action of the displacement operator on the vacuum state: |\alpha\rangle = D(\alpha) |0\rangle, with D(\alpha) = \exp(\alpha a^\dagger - \alpha^* a), and a^\dagger, a the ladder operators. This definition arises from the eigenvalue equation a |\alpha\rangle = \alpha |\alpha\rangle, highlighting their role as right eigenstates of the annihilation operator.^[19] In position space, the wavefunction of the coherent state is a Gaussian centered at the classical position x_{\rm cl} with a phase incorporating the classical momentum p_{\rm cl}:

\psi_\alpha(x) = \left( \frac{m\omega}{\pi \hbar} \right)^{1/4} \exp\left[ -\frac{m\omega}{2\hbar} (x - x_{\rm cl})^2 + i \frac{p_{\rm cl} (x - x_{\rm cl}/2)}{\hbar} + i \phi \right],

where x_{\rm cl}(t) = \sqrt{\frac{2\hbar}{m\omega}} |\alpha| \cos(\omega t + \phi) and p_{\rm cl}(t) = -\sqrt{2 m \omega \hbar} |\alpha| \sin(\omega t + \phi) mimic the classical solution for the oscillator. This form ensures the state achieves the minimum uncertainty \Delta x \Delta p = \hbar/2, with equal variances in position and momentum quadratures. Under time evolution, a coherent state remains coherent, evolving simply as |\alpha(t)\rangle = |\alpha e^{-i\omega t}\rangle, which corresponds to rigid rotation in phase space around the origin, preserving the Gaussian shape and uncertainty without distortion or spreading. This property underscores their quasi-classical behavior for the harmonic oscillator.^[19] Coherent states form an overcomplete basis, satisfying the resolution of the identity

\int \frac{d^2\alpha}{\pi} |\alpha\rangle \langle \alpha| = \hat{1},

allowing any state in the Hilbert space to be expanded in this basis, which is particularly useful in quantum optics and phase-space methods. The photon (or excitation) number distribution in a coherent state is Poissonian, with mean \langle n \rangle = |\alpha|^2 and variance equal to the mean, \Delta n^2 = |\alpha|^2, with relative fluctuations \Delta n / \langle n \rangle = 1 / \sqrt{\langle n \rangle} that become small in the classical limit of large |\alpha|, approaching the behavior of classical coherent light.^[19]

Squeezed States

Squeezed states represent a class of quantum states for the harmonic oscillator that achieve reduced uncertainty in one quadrature of the field at the expense of increased uncertainty in the conjugate quadrature, while still satisfying the Heisenberg uncertainty principle. These states generalize coherent states by applying a unitary squeeze operator to the vacuum or displaced vacuum. The squeeze operator is defined as S(\zeta) = \exp\left[ \frac{1}{2} (\zeta^* a^2 - \zeta (a^\dagger)^2) \right], where \zeta = r e^{i\theta} is a complex parameter with r denoting the squeeze parameter and \theta the squeeze angle, and a (a^\dagger) are the annihilation (creation) operators. A general squeezed coherent state is then given by |\zeta, \alpha\rangle = D(\alpha) S(\zeta) |0\rangle, where D(\alpha) = \exp(\alpha a^\dagger - \alpha^* a) is the displacement operator and |0\rangle is the vacuum state.^[20] The quadrature operators, which correspond to amplitude and phase quadratures in quantum optics, are X = \frac{a + a^\dagger}{\sqrt{2}} and P = \frac{a - a^\dagger}{i\sqrt{2}}. For a squeezed vacuum state with \theta = 0 and \alpha = 0, the uncertainties become \Delta X = e^{-r} \Delta X_{\text{vac}} and \Delta P = e^{r} \Delta P_{\text{vac}}, where \Delta X_{\text{vac}} = \Delta P_{\text{vac}} = \frac{1}{\sqrt{2}} for the vacuum state. This results in the product \Delta X \Delta P = \frac{1}{2}, saturating the Heisenberg limit \Delta X \Delta P \geq \frac{1}{2}. Squeezing is quantified by the minimum uncertainty being less than the vacuum value, i.e., \min(\Delta X, \Delta P) < \frac{1}{\sqrt{2}}, allowing for noise reduction in one observable below the standard quantum limit of coherent states.^[21] Squeezed states are generated experimentally through nonlinear optical processes, such as parametric down-conversion in a nonlinear crystal within an optical cavity, where a pump photon splits into signal and idler photons, producing correlated quadrature fluctuations. The first observation of squeezing via degenerate parametric down-conversion achieved noise reductions of 50% below the quantum limit. These states enable noise reduction beyond the limits achievable with coherent states (the unsqueezed limit at r = 0), with applications in precision measurements like gravitational wave detection, as explored in later sections.^[22]

Phase-Space Formulations

Phase-space formulations offer a powerful framework for visualizing and analyzing quantum states of the harmonic oscillator by mapping them onto classical-like phase-space distributions, revealing both quantum interference effects and classical correspondences. These representations, such as quasi-probability distributions, allow computation of expectation values via phase-space integrals analogous to classical statistical mechanics, while accommodating quantum features like non-commutativity through the parameter ℏ. The Wigner quasi-probability function stands as the most prominent example, providing a symmetric treatment of position and momentum. It is defined for a pure state with wave function ψ(x) as

W(x,p) = \frac{1}{\pi \hbar} \int_{-\infty}^{\infty} \psi^*(x+y) \psi(x-y) e^{2 i p y / \hbar} \, dy,

where the integral transforms the position-space density into a phase-space function whose marginals recover the position and momentum probability densities. Unlike classical probabilities, W(x,p) can assume negative values, signaling quantum non-classicality, and its integral over phase space equals unity. For the energy eigenstates |n⟩ of the quantum harmonic oscillator, the Wigner function adopts an explicit analytic form involving Laguerre polynomials, reflecting the oscillator's algebraic structure. In scaled variables where the classical Hamiltonian appears naturally, it reads

W_n(x,p) = \frac{(-1)^n}{\pi \hbar} \exp\left( -\frac{2 H_\mathrm{cl}}{\hbar \omega} \right) L_n \left( \frac{4 H_\mathrm{cl}}{\hbar \omega} \right),

with H_\mathrm{cl} = \frac{p^2}{2m} + \frac{1}{2} m \omega^2 x^2 the classical Hamiltonian, and L_n the nth Laguerre polynomial. These functions are azimuthally symmetric in phase space and, for n ≥ 1, exhibit regions of negativity due to the oscillatory nature of the Laguerre polynomials, quantifying deviations from classical behavior through non-positive quasi-probabilities.^[23]^[24] The ground state (n=0) yields a purely positive Gaussian distribution,

W_0(x,p) = \frac{1}{\pi \hbar} \exp\left( -\frac{2 H_\mathrm{cl}}{\hbar \omega} \right),

which matches the classical Boltzmann factor for the oscillator at infinite temperature in the high-energy limit but remains bounded and centered at the origin, illustrating minimal quantum spreading.^[23] Coherent states, which minimize uncertainty and follow classical trajectories under the oscillator Hamiltonian, have Wigner functions that are displaced Gaussians of the same form as the ground state but centered at the classical phase-space point (x_0, p_0) determined by the displacement parameter α, with |α|^2 giving the mean photon number. These distributions are always non-negative, preserving the classical-like positivity while orbiting the origin periodically with frequency ω. Other phase-space representations complement the Wigner function by addressing its negativity or sharpness. The Husimi Q-function, defined as Q(α) = (1/π) ⟨α| ρ |α⟩ where |α⟩ are coherent states and ρ the density operator, yields a positively valued, convolution-smoothed version of the Wigner function, useful for optical tomography despite added uncertainty. In contrast, the Glauber-Sudarshan P-representation expands ρ as an integral over coherent states with a distribution P(α) that can be highly singular or negative, facilitating calculations in quantum optics but requiring careful handling for non-classical states.

Multi-Dimensional Quantum Harmonic Oscillators

Isotropic N-Dimensional Case

The isotropic N-dimensional quantum harmonic oscillator models a particle confined by a quadratic potential that is identical in all spatial directions, characterized by the same angular frequency ω in each dimension. This generalization extends the one-dimensional case to higher dimensions while maintaining separability in Cartesian coordinates. The system's Hamiltonian takes the form of a sum over independent one-dimensional contributions, reflecting the absence of coupling between directions. The Hamiltonian is given by

\hat{H} = \sum_{i=1}^N \left( \frac{\hat{p}_i^2}{2m} + \frac{1}{2} m \omega^2 \hat{x}_i^2 \right) = \sum_{i=1}^N \hat{H}_i,

where \hat{H}_i = \frac{\hat{p}_i^2}{2m} + \frac{1}{2} m \omega^2 \hat{x}_i^2 is the one-dimensional Hamiltonian along the i-th coordinate axis, with \hat{p}_i and \hat{x}_i the momentum and position operators, respectively, and m the particle mass.^[25] Due to the separable structure, the time-independent Schrödinger equation factors into N independent one-dimensional equations, each solved using the standard harmonic oscillator approach. The corresponding energy eigenvalues are the sum of the individual one-dimensional energies:

E_{n_1, \dots, n_N} = \hbar \omega \left( \sum_{i=1}^N n_i + \frac{N}{2} \right),

where each n_i = 0, 1, 2, \dots is the quantum number for the i-th dimension.^[25] Introducing the total quantum number n = \sum_{i=1}^N n_i, the energies simplify to E_n = \hbar \omega (n + N/2), with n = 0, 1, 2, \dots. Each energy level E_n exhibits degeneracy, as multiple combinations of \{n_i\} yield the same n; the degeneracy factor is the binomial coefficient

g(n) = \binom{n + N - 1}{N - 1},

representing the number of distinct non-negative integer solutions to \sum n_i = n.^[26] This "accidental" degeneracy, beyond what the rotational symmetry alone would predict, stems from an underlying SU(N) algebraic structure governing the creation and annihilation operators across dimensions.^[26] The energy eigenfunctions are products of the one-dimensional harmonic oscillator wavefunctions:

\psi_{n_1, \dots, n_N}(x_1, \dots, x_N) = \prod_{i=1}^N \psi_{n_i}(x_i),

where each \psi_{n_i}(x_i) is the standard Hermite-Gaussian function for the i-th coordinate.^[25] The potential V(\mathbf{x}) = \frac{1}{2} m \omega^2 \sum_{i=1}^N x_i^2 = \frac{1}{2} m \omega^2 |\mathbf{x}|^2 is rotationally invariant under the orthogonal group SO(N), ensuring that the spectrum respects this symmetry while the full degeneracy reveals additional hidden symmetries.^[27]

Anisotropic and 3D Examples

In the anisotropic quantum harmonic oscillator, the frequencies differ along each spatial direction, leading to a Hamiltonian that does not possess full rotational symmetry. \hat{H} = \sum_{i=1}^{3} \left( \frac{\hat{p}_i^2}{2m} + \frac{1}{2} m \omega_i^2 \hat{x}_i^2 \right), where \omega_x, \omega_y, and \omega_z are the distinct angular frequencies, m is the particle mass, and \hat{p}_i, \hat{x}_i are the momentum and position operators in the i-th direction.^[28] This form separates completely in Cartesian coordinates, yielding energy eigenvalues E = \sum_{i=1}^{3} \hbar \omega_i \left( n_i + \frac{1}{2} \right), with non-negative integers n_x, n_y, n_z = 0, 1, 2, \dots.^[28] Unlike the isotropic case, the energy levels are generally non-degenerate unless the frequencies are commensurate (rationally related), in which case accidental degeneracies can arise from underlying symmetries.^[29] The corresponding eigenfunctions are products of one-dimensional harmonic oscillator wavefunctions in each direction, \psi_{n_x n_y n_z}(\mathbf{r}) = \psi_{n_x}(x) \psi_{n_y}(y) \psi_{n_z}(z), where each \psi_{n_i} is the standard 1D form involving Hermite polynomials.^[30] For the three-dimensional isotropic case, where \omega_x = \omega_y = \omega_z = \omega, the Hamiltonian simplifies to \hat{H} = \frac{\hat{p}^2}{2m} + \frac{1}{2} m \omega^2 r^2, exhibiting full SO(3) rotational invariance and allowing separation of variables in either Cartesian or spherical coordinates.^[31] In Cartesian coordinates, the energies follow the N=3 specialization of the isotropic spectrum, E = \hbar \omega (N/2 + n_x + n_y + n_z) with N=3, but spherical coordinates reveal the connection to angular momentum. In spherical coordinates (r, \theta, \phi), the Schrödinger equation separates into radial and angular parts, with the angular solutions being spherical harmonics Y_{l m}(\theta, \phi) labeled by orbital angular momentum quantum number l = 0, 1, 2, \dots and magnetic quantum number m = -l, \dots, l.^[31] The radial equation resembles that of a 1D oscillator but effective for l(l+1) centrifugal term, yielding associated Laguerre polynomial solutions with radial quantum number n_r = 0, 1, 2, \dots. The full quantum numbers are related by n = 2 n_r + l, where n = 0, 1, 2, \dots is the principal quantum number, and the energy is E_n = \hbar \omega \left( n + \frac{3}{2} \right). The eigenfunctions take the form \psi_{n l m}(r, \theta, \phi) = N_{n l} \, r^l \, e^{-m \omega r^2 / (2 \hbar)} \, L_{n_r}^{l + 1/2} \left( \frac{m \omega r^2}{\hbar} \right) Y_{l m}(\theta, \phi), where N_{n l} is a normalization constant and L denotes associated Laguerre polynomials. This structure couples the radial motion to the angular momentum \mathbf{L}^2 = \hbar^2 l(l+1) and L_z = \hbar m, mirroring central potentials like the hydrogen atom but with a quadratic confining potential instead of Coulombic.^[31] The ground state corresponds to n=0, l=0, m=0, a spherically symmetric s-state with wavefunction \psi_{000} \propto e^{-m \omega r^2 / (2 \hbar)} and energy E_0 = \frac{3}{2} \hbar \omega, fully non-degenerate. The first excited level has n=1, l=1, m = -1, 0, 1, forming a degenerate p-state triplet with energy E_1 = \frac{5}{2} \hbar \omega, where the three states transform as the l=1 irreducible representation under rotations, analogous to atomic p-orbitals.^[31] Higher levels exhibit further degeneracies, such as the n=2 manifold combining l=0 (s) and l=2 (d) states at E_2 = \frac{7}{2} \hbar \omega, with total degeneracy \frac{(n+1)(n+2)}{2}.

Applications in Physics

Phonons in Solid-State Lattices

In the study of lattice vibrations in crystalline solids, the harmonic approximation provides a foundational model by expanding the interatomic potential energy in a Taylor series around the equilibrium positions of the atoms, retaining only the linear and quadratic terms. The linear terms vanish at equilibrium, leaving the quadratic terms that describe small oscillations as a system of coupled harmonic oscillators, where the atoms interact via effective spring-like forces derived from second derivatives of the potential. This approximation is valid for temperatures well below the melting point, where anharmonic effects are minimal.^[32] To solve the dynamics of this coupled system, the equations of motion are diagonalized into independent normal modes through a Fourier transform, decoupling the oscillators into plane-wave-like vibrations propagating through the lattice. In the simplest case of a one-dimensional monatomic chain with nearest-neighbor interactions characterized by spring constant K, atomic mass m, and lattice spacing a, the dispersion relation for these normal modes is given by

\omega(k) = 2 \sqrt{\frac{K}{m}} \left| \sin\left( \frac{ka}{2} \right) \right|,

where k is the wavevector in the first Brillouin zone -\pi/a < k \leq \pi/a. This relation, derived under periodic boundary conditions (Born-von Kármán conditions), shows that the frequency vanishes at k = 0 (uniform translation) and reaches a maximum at the zone boundary.^[33] Quantizing these normal modes treats each as a quantum harmonic oscillator, with the vibrational quanta known as phonons—bosonic quasiparticles that carry energy and momentum through the lattice. The creation and annihilation operators b^\dagger_k and b_k for mode k satisfy the bosonic commutation relations [b_k, b^\dagger_{k'}] = \delta_{kk'}, and the corresponding Hamiltonian for the lattice vibrations becomes

\hat{H} = \sum_k \hbar \omega_k \left( b^\dagger_k b_k + \frac{1}{2} \right),

where the sum runs over all modes, and the zero-point contribution \frac{1}{2} \hbar \omega_k per mode reflects the ground-state energy of each oscillator.^[34]^[35] In three-dimensional solids, the dispersion relations exhibit acoustic branches where, for long wavelengths (k \to 0), the phonons behave like sound waves with a linear relation \omega = v |k|, v being the speed of sound; this is approximated in the Debye model by treating the lattice as a continuum of such modes up to a cutoff frequency. The model extends naturally to multi-dimensional anisotropic cases by considering vector displacements in the full lattice.^[34] Thermal properties of the lattice arise from the phonon excitations: the zero-point energy contributes a temperature-independent term \frac{1}{2} \sum_k \hbar \omega_k to the internal energy, while at high temperatures, the classical equipartition theorem assigns k_B T per degree of freedom to each mode, yielding a molar heat capacity of $3 N k_B (Dulong-Petit law) for N atoms in three dimensions, as each atom has three quadratic terms in kinetic and potential energy.^[34]^[35]

Molecular Vibrations and Spectroscopy

In the quantum mechanical treatment of diatomic molecules, the vibrational motion is approximated by a harmonic oscillator model, where the two atoms are connected by a spring-like force constant k. The effective mass for this oscillation is the reduced mass \mu = \frac{m_1 m_2}{m_1 + m_2}, with m_1 and m_2 being the atomic masses, leading to the angular frequency \omega = \sqrt{k / \mu}.^[36]^[7] The quantized energy levels are then given by E_v = \hbar \omega \left(v + \frac{1}{2}\right), where v = 0, 1, 2, \dots is the vibrational quantum number, establishing equally spaced levels that match experimental observations for low-lying states.^[36] This harmonic approximation holds well for small-amplitude vibrations near the equilibrium bond length but breaks down at higher energies due to anharmonicity from the true molecular potential, which is asymmetric and supports dissociation. The Morse potential, V(r) = D_e \left(1 - e^{-\alpha (r - r_e)}\right)^2, where D_e is the dissociation energy, \alpha relates to the curvature at equilibrium r_e, provides a more accurate description by incorporating cubic and higher-order terms as perturbations to the harmonic model.^[37] For low vibrational quantum numbers v, the harmonic oscillator remains a valid approximation, as anharmonic corrections shift levels by amounts proportional to v(v+1), which are small compared to \hbar \omega.^[38] For polyatomic molecules with N > 2 atoms, the vibrational degrees of freedom are described by $3N - 6 independent normal modes (or $3N - 5 for linear molecules), each treated as a quantum harmonic oscillator after transforming the Cartesian coordinates into a set of orthogonal normal coordinates that diagonalize the potential energy matrix.^[39] These normal modes represent collective motions where all atoms oscillate in phase with frequencies determined by the eigenvalues of the mass-weighted Hessian matrix, \omega_i = \sqrt{\lambda_i / \mu_i}, and the total vibrational energy is the sum over modes E = \sum_i \hbar \omega_i (v_i + 1/2).^[39] The $3N - 6 frequencies arise after subtracting three translational and three (or two) rotational degrees of freedom, enabling the analysis of complex spectra in terms of decoupled oscillators.^[40] Infrared (IR) spectroscopy probes these vibrational transitions, governed by selection rules derived from the harmonic oscillator matrix elements and the dipole approximation. For a fundamental transition, the vibrational quantum number changes by \Delta v = \pm 1 per mode, requiring a nonzero transition dipole moment \mu_{fi} = \langle f | \hat{\mu} | i \rangle, which necessitates a change in the molecular dipole moment during the vibration.^[41] Only modes that alter the dipole—such as asymmetric stretches in heteronuclear diatomics or specific bends in polyatomics like CO_2—are IR-active, while symmetric modes in homonuclear species are inactive.^[41] Rovibrational spectra of diatomic molecules reveal fine structure from coupled rotational and vibrational excitations, observed as bands in the IR with P, Q, and R branches corresponding to rotational changes \Delta J = -1, 0, +1, respectively, where J is the rotational quantum number.^[42] In the rigid rotor-harmonic oscillator approximation, the band origin marks the pure vibrational transition at \tilde{\nu}_0 = \omega / (2\pi c), with P-branch lines at lower wavenumbers (decreasing J), R-branch at higher (increasing J), and Q-branch (if allowed) at the origin for perpendicular transitions in polyatomics.^[42] This structure allows extraction of bond lengths and force constants from spacing analysis, with centrifugal distortion providing additional refinement.^[42]

Quantum Optics and Field Modes

In quantum optics, the electromagnetic field confined within a cavity is quantized by treating each spatial mode as an independent quantum harmonic oscillator. The Hamiltonian for the free field takes the form \hat{H} = \sum_k \hbar \omega_k \left( \hat{a}_k^\dagger \hat{a}_k + \frac{1}{2} \right), where \hat{a}_k^\dagger and \hat{a}_k are the creation and annihilation operators for photons in mode k with frequency \omega_k. This quantization of the cavity modes as harmonic oscillators resolves the classical ultraviolet catastrophe in blackbody radiation, leading to the Planck distribution for the thermal average photon number \langle n_k \rangle = 1/(e^{\hbar \omega_k / k_B T} - 1), which accurately describes the spectral energy density of blackbody radiation.^[43] The electric field operator is expressed as \hat{\mathbf{E}}(\mathbf{r}) \propto \sum_k \left( \frac{\hbar \omega_k}{2 \epsilon_0 V} \right)^{1/2} \left( \hat{a}_k \mathbf{u}_k(\mathbf{r}) + \hat{a}_k^\dagger \mathbf{u}_k^*(\mathbf{r}) \right), where \mathbf{u}_k(\mathbf{r}) are the mode functions normalized over the cavity volume V.^[44] This quantization scheme reveals the bosonic nature of photons and underpins phenomena like vacuum fluctuations, even in the absence of real photons.^[45] Photons emerge as quantized excitations of these field modes, with the number states |n_k\rangle for mode k representing n_k indistinguishable bosons obeying the harmonic oscillator energy levels E_n = \hbar \omega (n + 1/2).^[46] The ground state, or vacuum |0\rangle, exhibits zero-point fluctuations that contribute to measurable effects such as the Casimir force and spontaneous emission rates in cavities.^[44] These states form the Fock basis for describing light-matter interactions, where the harmonic oscillator analogy allows exact solutions for single-mode fields.^[45] A cornerstone application is the Jaynes-Cummings model, which describes the interaction between a two-level atom and a single-mode quantized field under the rotating-wave approximation, yielding the Hamiltonian \hat{H} = \hbar \omega \hat{a}^\dagger \hat{a} + \frac{\hbar \omega_0}{2} \hat{\sigma}_z + \hbar g (\hat{a}^\dagger \hat{\sigma}_- + \hat{a} \hat{\sigma}_+), where g is the coupling strength, \omega_0 the atomic transition frequency, and \hat{\sigma} the Pauli operators. For resonant conditions (\omega = \omega_0), this leads to vacuum Rabi oscillations, where the excitation energy oscillates between the atom and field at frequency $2g, observable in high-finesse cavities with Rydberg atoms.^[45] The model predicts collapses and revivals in the atomic inversion for coherent initial field states, highlighting quantum correlations absent in semiclassical treatments. Squeezed light, a nonclassical state of the field, is generated through nonlinear optical processes in media like parametric down-conversion crystals, where a pump photon splits into signal and idler modes, producing quadrature squeezing that reduces photon number noise below the vacuum level.^[47] In a degenerate optical parametric oscillator, the squeezing parameter r satisfies \langle (\Delta \hat{X})^2 \rangle = e^{-2r}/4 for one quadrature \hat{X}, enabling sub-shot-noise detection and improved precision in interferometry.^[47] This reduction in uncertainty, balanced by increased noise in the conjugate quadrature per the Heisenberg principle, has been demonstrated with up to 15 dB squeezing in continuous-wave experiments using potassium titanyl phosphate crystals. Cavity quantum electrodynamics (QED) leverages the harmonic oscillator model for single-mode fields in high-quality-factor resonators, such as Fabry-Pérot cavities, to achieve strong light-matter coupling where g > \kappa, \gamma (\kappa cavity decay, \gamma atomic decay).^[44] In this regime, the normal modes hybridize into polaritons, enabling coherent control of quantum states, as seen in microwave experiments with superconducting circuits mimicking atomic transitions.^[45] Cavity QED thus provides a testbed for fundamental quantum effects, including entanglement generation between atoms via photon exchange.^[44] This mode expansion of the electromagnetic field as an infinite collection of independent quantum harmonic oscillators extends naturally to quantum field theory, where relativistic quantum fields (scalar, vector, or spinor) are similarly decomposed into Fourier modes, each behaving as a harmonic oscillator. This framework allows the description of particles as excitations of these modes and facilitates the quantization of interacting fields, forming the basis for the standard model of particle physics.^[48]

Mathematical Extensions

Time-Dependent Solutions

The time-dependent Schrödinger equation governs the evolution of the wave function \psi(x, t) for the quantum harmonic oscillator:

i \hbar \frac{\partial \psi}{\partial t} = \hat{H} \psi,

where \hat{H} = \frac{\hat{p}^2}{2m} + \frac{1}{2} m \omega^2 \hat{x}^2 is the time-independent Hamiltonian. The quadratic form of \hat{H} ensures exact solvability, as the propagator can be constructed analytically using Gaussian integrals or operator methods, preserving the Gaussian structure of initial wave packets. For the undriven case, the energy eigenstates \psi_n(x) evolve simply by acquiring dynamical phase factors:

\psi_n(x, t) = \psi_n(x) \, e^{-i E_n t / \hbar},

with E_n = \hbar \omega (n + 1/2), resulting in no spatial distortion over time—only a global phase accumulation. This stationary evolution highlights the oscillator's role as a model for stable quantum systems, where superpositions of eigenstates maintain their form up to phases. In the driven case, the Hamiltonian becomes time-dependent: \hat{H}(t) = \hat{H}_0 - f(t) \hat{x}, where f(t) is an external force. Exact solutions are obtained using the Lewis-Riesenfeld dynamical invariant method, which constructs a time-independent invariant operator \hat{I}(t) satisfying i \hbar \partial_t \hat{I} = [\hat{I}, \hat{H}], allowing the wave function to be expressed as \psi(x, t) = \sum_n c_n \phi_n(x, t) e^{i \gamma_n(t)}, with \phi_n the instantaneous eigenstates of \hat{I} and \gamma_n a phase. Alternatively, the interaction picture or time-ordered exponentials yield the propagator, revealing displaced Gaussian wave packets that follow classical trajectories modulated by quantum spreads. The Ehrenfest theorem establishes the classical correspondence for expectation values, yielding

\frac{d \langle \hat{x} \rangle}{dt} = \frac{\langle \hat{p} \rangle}{m}, \quad \frac{d \langle \hat{p} \rangle}{dt} = -m \omega^2 \langle \hat{x} \rangle + \langle f(t) \rangle,

which mirror the Newtonian equations for a driven classical oscillator, valid exactly for quadratic Hamiltonians. Higher moments, such as variances, evolve independently of the means, underscoring the separability of center-of-mass and relative motion. For periodic driving, where f(t + T) = f(t), Floquet theory applies, decomposing solutions into quasi-periodic forms \psi(x, t) = e^{-i \epsilon t / \hbar} \varphi(x, t) with \varphi(x, t + T) = \varphi(x, t) and \epsilon the quasi-energy. This framework reveals stable Floquet modes for the oscillator, analogous to Mathieu functions in the classical limit, enabling analysis of resonance and stability in periodically modulated systems.

Supersymmetric Quantum Mechanics

Supersymmetric quantum mechanics (SUSY QM) reformulates the quantum harmonic oscillator in terms of bosonic and fermionic degrees of freedom, revealing an underlying algebraic structure that extends the familiar ladder operator method. In this framework, the Hilbert space is a tensor product of the bosonic Fock space and a fermionic two-dimensional space spanned by states with zero or one fermion, allowing the construction of supercharges that map between these sectors. The supercharges are defined as Q = \sqrt{2 \hbar \omega} \, a \, \psi^\dagger and Q^\dagger = \sqrt{2 \hbar \omega} \, a^\dagger \, \psi, where a and a^\dagger are the bosonic annihilation and creation operators satisfying [a, a^\dagger] = 1, and \psi, \psi^\dagger are fermionic operators obeying the Clifford algebra \{ \psi, \psi^\dagger \} = 1. These satisfy the supersymmetry algebra Q^2 = (Q^\dagger)^2 = 0 and \{ Q, Q^\dagger \} = 2 \hbar \omega (a^\dagger a + \psi^\dagger \psi), where the full Hamiltonian is H = \frac{1}{2} \{ Q, Q^\dagger \} + \frac{1}{2} \hbar \omega = \hbar \omega \left( a^\dagger a + \psi^\dagger \psi + \frac{1}{2} \right).^[49] The spectrum of H consists of doubly degenerate levels for excited states, with energies E_n = \hbar \omega \left( n + \frac{1}{2} \right) for n \geq 1, corresponding to states |n, 0\rangle (bosonic) and |n-1, 1\rangle (fermionic), while the ground state |0, 0\rangle at E_0 = \frac{1}{2} \hbar \omega is uniquely bosonic and annihilated by both Q and Q^\dagger. This non-degeneracy of the ground state implies unbroken supersymmetry, as the supercharges cannot mix it with a fermionic partner. The Witten index, defined as \Delta = \mathrm{Tr} \left[ (-1)^F e^{-\beta H} \right] where F = \psi^\dagger \psi is the fermion number operator, evaluates to 1 for the harmonic oscillator, reflecting the excess of one bosonic ground state over fermionic ones and serving as a topological invariant independent of \beta. In the position representation, SUSY QM factorizes the Hamiltonian into partner operators, with the full H = \begin{pmatrix} H_- & 0 \\ 0 & H_+ \end{pmatrix} where H_- = A^\dagger A and H_+ = A A^\dagger, and A = \frac{d}{dx} + W(x) with superpotential W(x) = \frac{m \omega}{\hbar} x. The partner potentials are V_-(x) = \frac{\hbar^2}{2m} \left[ W^2(x) - W'(x) \right] and V_+(x) = \frac{\hbar^2}{2m} \left[ W^2(x) + W'(x) \right], which for the harmonic oscillator yield identical quadratic forms \frac{1}{2} m \omega^2 x^2 up to an energy shift of \hbar \omega, ensuring isospectrality except for the shared ground state.^[50] The harmonic oscillator exemplifies shape-invariant potentials in SUSY QM, where the partner potential satisfies V_+(x; a_0) = V_-(x; a_1) + R(a_0) with parameters shifted by a_1 = f(a_0) and remainder R(a_0) independent of x; here, the "parameter" is the constant term, leading to recursive generation of the energy spectrum E_n = E_0 + \sum_{k=0}^{n-1} R(a_k) = \hbar \omega \left( n + \frac{1}{2} \right). This property, first systematically exploited for solvable models, allows exact algebraic construction of eigenstates and highlights the oscillator's role in generating spectra for related potentials like the radial oscillator. Applications of this SUSY structure provide insights into unbroken supersymmetry, where the unique vacuum demonstrates stability without spontaneous breaking, contrasting with models exhibiting degeneracy. It also facilitates exact solutions for other shape-invariant systems, such as the hydrogen atom or Morse potential, by iterating the factorization method starting from the oscillator-like seed.

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