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Free particle

In physics, a free particle is a model for a particle subject to no external forces or potentials, enabling it to propagate indefinitely without acceleration or confinement.^[1] In classical mechanics, a free particle obeys Newton's first law, maintaining constant velocity and linear motion in a field-free space, with its kinetic energy given by E = \frac{p^2}{2m}, where p is momentum and m is mass.^[1]^[2] This idealized scenario serves as a foundational reference for understanding more complex dynamics under forces. In quantum mechanics, the free particle is described by the time-independent Schrödinger equation -\frac{\hbar^2}{2m} \frac{d^2 \psi(x)}{dx^2} = E \psi(x), where \psi(x) is the wave function, \hbar is the reduced Planck's constant, and E is energy.^[1] The solutions are plane waves of the form \psi_{\pm}(x) = A_{\pm} e^{\pm i k x}, with wave number k = \sqrt{2mE}/\hbar, corresponding to definite momenta p = \pm \hbar k and continuous energy spectra due to the absence of boundaries.^[1]^[3] These states exhibit uniform probability density over all space, illustrating the Heisenberg uncertainty principle: precise momentum knowledge implies complete positional delocalization.^[1] In relativistic quantum mechanics, free particles are described by wave equations such as the Klein–Gordon equation for spin-0 particles and the Dirac equation for spin-1/2 particles.^[4] The free particle model is crucial for deriving approximations in scattering theory, solid-state physics, and understanding wave propagation in unbounded systems.^[5]^[6]

Classical Mechanics

Definition and Kinematics

In classical mechanics, a free particle is defined as a point mass that experiences no net external force acting upon it. According to Newton's first law of motion, such a particle will continue in a state of rest or uniform motion in a straight line at constant velocity unless compelled to change its state by forces impressed upon it. This idealization assumes an absence of interactions with other bodies, fields, or constraints, allowing the particle's trajectory to be purely inertial. The kinematics of a free particle are straightforward and deterministic, governed solely by its initial conditions. The position of the particle as a function of time is given by

\mathbf{r}(t) = \mathbf{r}_0 + \mathbf{v} t,

where \mathbf{r}_0 is the initial position vector at t = 0, and \mathbf{v} is the constant velocity vector, independent of time. This linear trajectory implies zero acceleration (\mathbf{a} = d\mathbf{v}/dt = 0), and the speed |\mathbf{v}| remains unchanged. In one dimension, for simplicity, this reduces to x(t) = x_0 + v t, highlighting the uniform progression along the path. This behavior is fundamentally tied to the concept of inertial reference frames, where the laws of motion hold without fictitious forces. In such frames, the free particle's motion exhibits Galilean invariance: its velocity appears constant regardless of the frame's uniform translation relative to another inertial frame. This invariance underpins the relativity of uniform motion in non-relativistic physics, distinguishing free particles from those under acceleration. Examples of free particle motion include the idealized trajectory of a spacecraft far from gravitational influences in interplanetary space or an air molecule in a perfect vacuum, where collisions are negligible. In contrast, real particles are often constrained by potentials, such as gravitational or electromagnetic fields, leading to curved paths—unlike the unbound straight-line propagation of the truly free case.

Dynamics and Conservation Laws

The dynamics of a classical free particle are governed by the absence of external forces, leading directly to Newton's first law of motion, which states that a particle remains at rest or in uniform rectilinear motion unless acted upon by a net external force.^[7] This law establishes that the acceleration of the particle is zero, \mathbf{a} = 0, implying constant velocity \mathbf{v}.^[8] Due to translational symmetry in space—the invariance of the system's laws under spatial displacements—the linear momentum \mathbf{p} = m \mathbf{v} of the free particle is conserved, remaining constant throughout its motion.^[9] This conservation arises from Noether's theorem, which links continuous symmetries of the action to conserved quantities; for a free particle, the symmetry under infinitesimal translations \delta \mathbf{x} = \epsilon yields the continuity equation \frac{d\mathbf{p}}{dt} = 0.^[9] Similarly, time-translation invariance implies the conservation of total energy. For a free particle, the kinetic energy E = \frac{1}{2} m v^2 is the sole contribution and remains constant, as there are no non-conservative forces to alter it.^[10] This follows from the time-independence of the Lagrangian, ensuring \frac{dE}{dt} = 0 via the Euler-Lagrange formalism.^[11] In the Lagrangian formulation, the free particle's Lagrangian is L = \frac{1}{2} m \dot{x}^2 (in one dimension for simplicity), where the kinetic energy serves as the Lagrangian since potential energy vanishes.^[12] Applying the Euler-Lagrange equation \frac{d}{dt} \left( \frac{\partial L}{\partial \dot{x}} \right) - \frac{\partial L}{\partial x} = 0 gives \frac{d}{dt} (m \dot{x}) = 0, confirming constant velocity \dot{x} = v.^[13] The Hamiltonian formulation provides an equivalent description, with the Hamiltonian H = \frac{p^2}{2m} expressed in terms of momentum p = m \dot{x}.^[14] Hamilton's equations, \dot{x} = \frac{\partial H}{\partial p} = \frac{p}{m} and \dot{p} = -\frac{\partial H}{\partial x} = 0, reproduce the constant velocity and conserved momentum, while the time-independence of H underscores energy conservation.^[15]

Non-Relativistic Quantum Mechanics

Schrödinger Equation Formulation

In 1926, Erwin Schrödinger introduced the wave equation as part of his formulation of wave mechanics, with the free particle representing the simplest case where no external potential acts on the system. The derivation of the Schrödinger equation for a free particle proceeds from the classical Hamiltonian via canonical quantization rules. The classical Hamiltonian for a free particle of mass m is H = \frac{p^2}{2m}, where p is the momentum. In quantum mechanics, position x and momentum p are promoted to operators satisfying the commutation relation [x, p] = i\hbar, with p = -i\hbar \nabla in the position representation. Substituting into the quantum Hamiltonian yields the operator \hat{H} = -\frac{\hbar^2}{2m} \nabla^2. The time evolution of the wave function \psi(\mathbf{r}, t) follows from the correspondence principle, where the classical Poisson bracket is replaced by the commutator, leading to the time-dependent Schrödinger equation:

i\hbar \frac{\partial \psi}{\partial t} = -\frac{\hbar^2}{2m} \nabla^2 \psi.

^[16] For stationary states, where the wave function separates as \psi(\mathbf{r}, t) = \phi(\mathbf{r}) e^{-iEt/\hbar}, the time-dependent equation reduces to the time-independent form:

-\frac{\hbar^2}{2m} \nabla^2 \phi = E \phi,

with E as the energy eigenvalue. This eigenvalue equation describes the spatial behavior of the particle in energy eigenstates.^[16] The free particle exists in infinite space with zero potential V = 0 everywhere, imposing no confining boundary conditions. Consequently, the domain of the wave function extends over all space, \mathbf{r} \in \mathbb{R}^3, without restrictions on the wave function at finite boundaries.^[17] The probabilistic interpretation, proposed by Max Born in 1926, assigns |\psi(\mathbf{r}, t)|^2 as the probability density for finding the particle at position \mathbf{r} at time t. For the equation to yield physically meaningful probabilities, the wave function must satisfy the normalization condition \int |\psi|^2 d^3\mathbf{r} = 1. However, in infinite space, exact energy eigenstates are plane waves that extend indefinitely, rendering them non-normalizable as their integral diverges; practical treatments often employ wave packets or impose periodic boundary conditions in a large but finite volume to approximate normalization.^[17]

Plane Wave Solutions

In non-relativistic quantum mechanics, the plane wave solutions to the free particle Schrödinger equation take the form

\psi(x,t) = A \exp\left[i (k x - \omega t)\right],

where A is a constant amplitude, k is the wave number, and \omega is the angular frequency. These functions satisfy the time-dependent Schrödinger equation i \hbar \frac{\partial \psi}{\partial t} = -\frac{\hbar^2}{2m} \frac{\partial^2 \psi}{\partial x^2} for zero potential. Substituting the plane wave into the Schrödinger equation yields the dispersion relation \omega = \frac{\hbar k^2}{2m}, linking the frequency to the wave number and particle mass m. This relation follows from the eigensolution property of the equation. The plane waves serve as eigenfunctions of the momentum operator \hat{p} = -i \hbar \frac{\partial}{\partial x}, with eigenvalue p = \hbar k. The associated energy eigenvalue is E = \hbar \omega = \frac{p^2}{2m}, reflecting the classical kinetic energy in quantum form. Plane waves extend infinitely in space and thus are not square-integrable, preventing normalization in the usual L^2 sense over infinite domains. Instead, they employ Dirac delta function normalization, where the position representation satisfies \int_{-\infty}^{\infty} \psi_p^*(x) \psi_{p'}(x) \, dx = \delta(p - p'), ensuring orthogonality for different momenta. By the superposition principle, any general solution to the free particle Schrödinger equation can be expressed as an integral over plane waves:

\psi(x,t) = \frac{1}{\sqrt{2\pi}} \int_{-\infty}^{\infty} \tilde{\psi}(k) \exp\left[i (k x - \omega(k) t)\right] \, dk,

with \tilde{\psi}(k) as the Fourier amplitude and \omega(k) = \frac{\hbar k^2}{2m}. This Fourier decomposition expands arbitrary initial wave functions in the continuous momentum basis. Physically, plane waves describe delocalized states with precise momentum p = \hbar k but infinite uncertainty in position, embodying the trade-off in the Heisenberg uncertainty principle \Delta x \Delta p \geq \hbar/2. Such states are idealized, representing particles with uniform probability density across all space.

Wave Packets and Superpositions

In quantum mechanics, plane waves provide exact solutions to the free particle Schrödinger equation but are delocalized over all space, failing to represent a particle confined to a finite region. To model a localized particle with an approximate position and momentum, the wave function is constructed as a superposition of plane waves with a continuous range of wave numbers centered around a mean value. This approach, introduced by Schrödinger as "wave groups," allows the probability density to be concentrated in a small spatial region while satisfying the superposition principle of the Schrödinger equation.^[18] The general form of such a wave packet in one dimension is given by the Fourier integral representation

\psi(x,t) = \int_{-\infty}^{\infty} \mathrm{d}k \, \phi(k) \, \exp\left[i \left( k x - \omega(k) t \right)\right],

where \phi(k) is the amplitude in momentum space, serving as the Fourier transform of the initial wave function, and \omega(k) = \hbar k^2 / (2m) is the angular frequency for a free particle of mass m. The function \phi(k) is typically chosen to be peaked around some central wave number k_0, ensuring a spread in momenta \Delta k that corresponds to a spatial localization via the uncertainty principle \Delta x \Delta p \gtrsim \hbar/2. This representation normalizes the wave function such that \int |\psi(x,t)|^2 \mathrm{d}x = 1, provided \int |\phi(k)|^2 \mathrm{d}k = 1. A common example is the Gaussian wave packet, which minimizes the position-momentum uncertainty and provides an analytically tractable form. At t=0, it takes the form

\psi(x,0) = \left( \frac{1}{2\pi \sigma^2} \right)^{1/4} \exp\left[ -\frac{x^2}{4\sigma^2} + i k_0 x \right],

where \sigma is the initial spatial width. The corresponding \phi(k) is also Gaussian, centered at k_0 with width $1/(2\sigma). This initial state is localized around the mean position \langle x \rangle = 0 (or shifted accordingly), with mean momentum \langle p \rangle = \hbar k_0, as computed from the expectation values \langle x \rangle = \int \psi^* x \psi \, \mathrm{d}x and \langle p \rangle = -i\hbar \int \psi^* \frac{\partial \psi}{\partial x} \, \mathrm{d}x. Such wave packets bridge classical and quantum descriptions by having their center of probability density follow a classical trajectory under the Ehrenfest theorem, \frac{\mathrm{d} \langle x \rangle}{\mathrm{d}t} = \langle p \rangle / m and \frac{\mathrm{d} \langle p \rangle}{\mathrm{d}t} = -\left\langle \frac{\partial V}{\partial x} \right\rangle, which for a free particle (V=0) yields uniform motion. However, since the underlying plane waves extend infinitely, constructing normalizable wave packets in unbounded space demands careful choice of \phi(k) with finite support or rapid decay to ensure convergence of the integral and proper normalization.

Group and Phase Velocities

In non-relativistic quantum mechanics, the free particle is described by plane wave solutions to the Schrödinger equation, which exhibit a dispersion relation \omega(k) = \frac{\hbar k^2}{2m}, where \omega is the angular frequency, k is the wave number, \hbar is the reduced Planck's constant, and m is the particle mass.^[19] This relation arises directly from substituting the plane wave form \psi(x,t) = e^{i(kx - \omega t)} into the time-dependent Schrödinger equation for zero potential, i\hbar \frac{\partial \psi}{\partial t} = -\frac{\hbar^2}{2m} \frac{\partial^2 \psi}{\partial x^2}.^[19] The phase velocity v_p of these waves is defined as the speed at which a constant phase point propagates, given by v_p = \frac{\omega}{k} = \frac{\hbar k}{2m}.^[19] This velocity depends linearly on k, meaning it varies for different wavelength components and does not correspond to the motion of a classical particle. In contrast, the group velocity v_g, which represents the propagation speed of the overall wave envelope, is obtained by differentiating the dispersion relation: v_g = \frac{d\omega}{dk} = \frac{\hbar k}{m}.^[19] Since the momentum p = \hbar k, this simplifies to v_g = \frac{p}{m}, matching the classical velocity of a free particle with the same momentum.^[19] Physically, the phase velocity characterizes the motion of individual plane wave components, while the group velocity describes the velocity of the wave packet's envelope, which localizes the particle's probable position and thus aligns with the classical particle trajectory.^[19] For a wave packet constructed as a superposition of plane waves with a narrow distribution of k values around a central k_0, the group velocity approximates the classical velocity v = \frac{\hbar k_0}{m}, providing a clear quantum analog to particle motion.^[19]

Wave Packet Spreading and Uncertainty

In quantum mechanics, the wave function of a free particle, initially localized as a wave packet, undergoes temporal evolution governed by the Schrödinger equation, leading to a gradual broadening of the packet even in the absence of any potential. This spreading arises from the dispersion in the momentum components of the superposition, where plane waves with different wavelengths propagate at different velocities. For a Gaussian wave packet, which achieves the minimum uncertainty in position and momentum, the time-dependent width \sigma(t) is given by

\sigma(t)^2 = \sigma(0)^2 + \left( \frac{\hbar t}{2 m \sigma(0)} \right)^2,

where \sigma(0) is the initial width, m is the particle mass, \hbar is the reduced Planck's constant, and t is time.^[20]^[21] The derivation of this formula proceeds from the time evolution of the position variance, defined as \langle x^2 \rangle - \langle x \rangle^2. For a free particle, the expectation value \langle x \rangle moves with constant group velocity, but the second moment \langle x^2 \rangle increases due to the spread in velocities from the momentum distribution. Using the Fourier transform of the initial Gaussian wave function and applying the time-dependent phase factors e^{-i E(k) t / \hbar} with E(k) = \hbar^2 k^2 / (2m), the evolved wave function remains Gaussian, allowing exact computation of the moments via Gaussian integrals. This yields the quadratic growth in variance, reflecting the dispersive nature of the non-relativistic dispersion relation \omega(k) \propto k^2.^[20]^[22] This spreading directly illustrates the Heisenberg uncertainty principle, \Delta x \Delta p \geq \hbar / 2, where the initial minimum-uncertainty Gaussian satisfies \Delta x(0) \Delta p = \hbar / 2 with \Delta p = \hbar / (2 \Delta x(0)). As time progresses, the position uncertainty \Delta x(t) \approx \sigma(t) increases while \Delta p remains constant, maintaining the product above the minimum but demonstrating how the fixed momentum spread \Delta p leads to differential velocities \Delta v = \Delta p / m, causing the packet to broaden. A narrower initial localization (smaller \sigma(0)) implies a larger \Delta p, resulting in faster spreading, underscoring the principle's dynamic implications.^[21]^[22] Physically, this phenomenon implies that quantum particles exhibit diffusive-like behavior, delocalizing over time unlike classical point particles that maintain fixed trajectories without forces. For long times, t \gg 2 m \sigma(0)^2 / \hbar, the width grows linearly as \sigma(t) \approx \hbar t / (2 m \sigma(0)), with the rate inversely proportional to mass and initial width, making it pronounced for light particles like electrons but negligible for macroscopic objects due to the small scale of \hbar. This spreading highlights the inherently probabilistic nature of quantum propagation for free particles.^[21]^[20]

Relativistic Quantum Mechanics

Klein-Gordon Equation

The Klein-Gordon equation serves as the fundamental relativistic wave equation for free particles of spin zero, extending the principles of quantum mechanics to incorporate special relativity. It arises from the quantization of the classical relativistic energy-momentum relation E^2 = p^2 c^2 + m^2 c^4, where E is the energy, p the momentum magnitude, m the rest mass, and c the speed of light. By promoting E to the operator i \hbar \partial_t and p to -i \hbar \nabla, and squaring to obtain a second-order equation, the result is the Klein-Gordon equation in its time-dependent form:

\left( \frac{1}{c^2} \frac{\partial^2}{\partial t^2} - \nabla^2 + \frac{m^2 c^2}{\hbar^2} \right) \psi(\mathbf{x}, t) = 0,

or, in covariant notation using the d'Alembertian operator \square = \partial^\mu \partial_\mu,

(\square + \frac{m^2 c^2}{\hbar^2}) \psi(x) = 0.

This equation was independently derived by Oskar Klein and Walter Gordon in 1926, predating Paul Dirac's work on spin-1/2 particles, as part of early efforts to reconcile quantum mechanics with relativity. Multiple physicists, including Vladimir Fock, contributed similar derivations that year, reflecting the equation's rapid emergence in the literature. The formulation ensures Lorentz invariance, treating the wave function \psi as a scalar field describing the particle's probability amplitude. Plane wave solutions to the Klein-Gordon equation take the form \psi(\mathbf{x}, t) \propto \exp\left[-i (E t - \mathbf{p} \cdot \mathbf{x})/\hbar \right], where the dispersion relation is E = \pm \sqrt{p^2 c^2 + m^2 c^4}. These solutions include both positive-energy states (corresponding to particles) and negative-energy states (initially interpreted as antiparticles or problematic excitations). Superpositions of such plane waves can form wave packets that propagate relativistically, with the group velocity v_g = \partial E / \partial p = p c^2 / E matching the classical particle velocity for positive-energy components. Despite its formal successes, the Klein-Gordon equation encounters significant challenges in single-particle quantum mechanics. The conserved four-current, derived from Noether's theorem for the global U(1) phase invariance, yields a charge density \rho = (i \hbar / 2 m c^2) (\psi^* \partial_t \psi - \psi \partial_t \psi^*) that is not positive definite, allowing negative values due to interference between positive- and negative-energy components.^[23] This undermines a probabilistic interpretation, as the "probability density" can become negative or fail to normalize properly. In the non-relativistic limit (low momenta, p \ll m c), the equation reduces formally to the Schrödinger equation for positive-energy projections, but the density issues persist for low-energy states, where relativistic corrections introduce oscillations and non-positivity not present in the non-relativistic case. These limitations, recognized soon after its proposal, highlighted the need for a second-quantized field-theoretic interpretation, where the equation describes a quantum field with particle and antiparticle creation/annihilation rather than a single-particle wave function.

Dirac Equation Solutions

The Dirac equation provides a relativistic wave equation for free particles with spin-1/2, such as electrons, incorporating both quantum mechanics and special relativity. It is expressed in Hamiltonian form as

i \hbar \frac{\partial \psi}{\partial t} = c \boldsymbol{\alpha} \cdot \mathbf{p} \psi + \beta m c^2 \psi,

where \psi is a four-component spinor wave function, \mathbf{p} = -i \hbar \boldsymbol{\nabla} is the momentum operator, m is the particle mass, c is the speed of light, \hbar is the reduced Planck's constant, and \boldsymbol{\alpha} = (\alpha_x, \alpha_y, \alpha_z) along with \beta are 4×4 Hermitian matrices obeying the anticommutation relations \{\alpha_i, \alpha_j\} = 2\delta_{ij}, \{\alpha_i, \beta\} = 0, and \beta^2 = 1.^[24] These matrices ensure the equation is linear and first-order in both space and time derivatives, addressing limitations of prior relativistic formulations like the Klein-Gordon equation, which treated particles as scalars without spin and led to issues with probability interpretation.^[24] Plane wave solutions to the Dirac equation for free particles are of the form \psi(\mathbf{x}, t) = u(\mathbf{p}) \exp\left[-i (E t - \mathbf{p} \cdot \mathbf{x})/\hbar\right] for positive energy states, where u(\mathbf{p}) is a four-component spinor normalized such that \bar{u} u = 2m, and \bar{u} = u^\dagger \beta.^[25] For negative energy states, the solutions are \psi(\mathbf{x}, t) = v(\mathbf{p}) \exp\left[i (E t - \mathbf{p} \cdot \mathbf{x})/\hbar\right], with v(\mathbf{p}) satisfying \bar{v} v = -2m and interpreted as describing antiparticles, such as positrons, in the Dirac sea picture.^[25] Substituting these into the Dirac equation yields (\slash{p} - m) u(\mathbf{p}) = 0 for positive energy and (\slash{p} + m) v(\mathbf{p}) = 0 for negative energy, where \slash{p} = \gamma^\mu p_\mu in covariant notation with Dirac matrices \gamma^\mu.^[25] The energy spectrum from these solutions is E = \pm \sqrt{p^2 c^2 + m^2 c^4}, where p = |\mathbf{p}|, providing both positive and negative branches that match the relativistic energy-momentum relation.^[25] This spectrum resolves the interpretive challenges of negative energy solutions in the Klein-Gordon equation—such as the Klein paradox involving apparent pair production—by reinterpreting them as holes in a filled negative energy sea, corresponding to antiparticles with positive energy.^[26] In the massless limit (m \to 0), the Dirac equation reduces to the Weyl equation, with solutions exhibiting two helicity states: particles with spin projection along the momentum direction (+1/2) and opposite (-1/2), reflecting left- and right-handed chiralities that coincide with helicity.^[27] Formulated by Paul Dirac in 1928, the equation predicted the existence of antimatter through its negative energy solutions, a prediction experimentally confirmed by the discovery of the positron by Carl D. Anderson in 1932 using cloud chamber observations of cosmic rays.^[24]^[28]

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